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Fluctuation-dissipation dynamics of cosmological scalar fields

Citation for published version:Bartrum, S, Berera, A & Rosa, JG 2015, 'Fluctuation-dissipation dynamics of cosmological scalar fields',Physical Review D - Particles, Fields, Gravitation and Cosmology, vol. 91, no. 8, 083540.https://doi.org/10.1103/PhysRevD.91.083540

Digital Object Identifier (DOI):10.1103/PhysRevD.91.083540

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Fluctuation-dissipation dynamics of cosmological scalar fields

Sam Bartrum,1,* Arjun Berera,1,† and João G. Rosa2,‡,§1SUPA, School of Physics and Astronomy, University of Edinburgh, Edinburgh EH9 3JZ, United Kingdom

2Departamento de Física da Universidade de Aveiro and CIDMA,Campus de Santiago, 3810-183 Aveiro, Portugal

(Received 8 January 2015; published 29 April 2015)

We show that dissipative effects have a significant impact on the evolution of cosmological scalarfields, leading to friction, entropy production and field fluctuations. We explicitly compute the dissipationcoefficient for different scalar fields within the standard model and some of its most widely consideredextensions, in different parametric regimes. We describe the generic consequences of fluctuation-dissipation dynamics in the postinflationary universe, focusing in particular on friction and particleproduction, and analyze in detail two important effects. First, we show that dissipative friction delays theprocess of spontaneous symmetry breaking and may even damp the motion of a Higgs field sufficiently toinduce a late period of warm inflation. Along with dissipative entropy production, this may parametricallydilute the abundance of dangerous thermal relics. Second, we show that dissipation can generate theobserved baryon asymmetry without symmetry restoration, and we develop in detail a model of dissipativeleptogenesis. We further show that this generically leads to characteristic baryon isocurvature perturbationsthat can be tested with cosmic microwave background observations. This work provides a fundamentalframework to go beyond the leading thermal equilibrium semiclassical approximation in addressingfundamental problems in modern cosmology.

DOI: 10.1103/PhysRevD.91.083540 PACS numbers: 98.80.-k, 11.30.Fs, 11.30.Qc, 98.80.Cq

I. INTRODUCTION

Scalar fields play a major role in modern cosmologicaltheories. Depending on the balance between the kinetic,potential and gradient energy stored in these fields, they canmimic fluids with distinct equations of state and so havebeen proposed as leading candidates to describe the earlyphase of inflationary expansion, as well as dark matter anddark energy.Scalar fields are also a key ingredient in modern particle

physics theories and the recent discovery of what is nowwidely accepted to be the electroweak Higgs boson at theLHC puts the existence of fundamental scalar degrees offreedom on firm experimental ground. Indeed, these areubiquitous in extensions of the standard model (SM) ofparticle physics, such as grand unified theories (GUT),supersymmetric theories or extradimensional scenarios,namely within the context of string/M-theory compactifi-cations. The study of the cosmological dynamics of scalarfields, both at the classical and quantum levels, is thus ofcrucial importance to understand the early history of ourUniverse.One of the most prominent roles of scalar fields is the

phenomenon of spontaneous symmetry breaking in

fundamental gauge theories, where vector bosons andfermions acquire mass through the Bose condensation ofa scalar field. This process of spontaneous symmetrybreaking sees an initial symmetric state go to a state ofbroken symmetry, all due to the change of a singleparameter, the vacuum expectation value (vev) of a scalarfield hϕi. The electroweak Higgs mechanism is the bestknown example of this simple idea, which is also expectedto apply to the spontaneous breaking of higher-rank gaugesymmetry groups that extend the SM at high-energy scales.The significance of spontaneous symmetry breaking

for cosmology was pointed out several decades ago byKirzhnits and Linde [1]. They observed that this behaviorwould be a feature of quantum field theories at finitetemperature, whereby at very high temperatures the vev ofthe scalar field would be a single value that restoressymmetry and then, at some specific critical temperature,this vev would change and lead to a phase of brokensymmetry. The description of this process fits well withinthe Landau theory of phase transitions. These two simpleideas of spontaneous symmetry breaking and its realizationin finite temperature quantum field theory as a phasetransition have been the foundation for cosmological phasetransitions [2,3]. Such behavior has since been applied tonumerous areas in cosmology including inflation, defects,baryogenesis [4], and cosmic magnetic fields [5].The study of cosmological phase transitions has so far

been centered primarily on their equilibrium properties.In particular, most of the interest has gone into studyingthe particle physics features in the symmetric and broken

*s.bartrum@sms.ed.ac.uk†ab@ph.ed.ac.uk‡Also at Departamento de Física e Astronomia, Faculdade de

Ciências da Universidade do Porto, Rua do Campo Alegre 687,4169-007 Porto, Portugal.

§joao.rosa@ua.pt

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phases. The dynamics that induces the change from onephase to the other is, however, also a necessary componentof this entire process. This change will involve the motionof the order parameter hϕi from the symmetry-restored tobroken phase. Since this scalar order parameter is theexpectation value of a quantum field, which in generalinteracts with other fields that comprise the radiation bath,its evolution between the different phases will genericallyinvolve energy exchange. Due to the tendency for theequipartition of energy in dynamical systems, this appearsprimarily as energy exchange between the single dynamicalscalar degree of freedom and the many degrees of freedomcomprising the heat bath. This thus results in dissipation ofthe scalar field’s energy into the ambient radiation fluid.The order parameter experiences another effect when

immersed in the radiation bath. All the random interactionsof the bath constituents with the single order parameterwill slightly push the scalar field around in all differentdirections, thus inducing fluctuations. These two processesof dissipation and fluctuations of the order parameterare intrinsically related to each other by the underlyingdynamical quantum mechanical equations [6]. This is thebasis of fluctuation-dissipation theorems and it is appli-cable to the dynamics of cosmological phase transitionsjust as it is to any phase transitions or out-of-equlibriumsituation in condensed matter systems [7–10].As such, wherever a cosmological phase transition is

present, fluctuation-dissipation dynamics will be presenthand in hand during the out-of-equilibrium transitionperiod between the two equilibrium phases. This phasetransition dynamics will add three new features to theequilibrium description. First, the background evolutionof the scalar order parameter will affect the expansionbehavior of the Universe. Second, there will be particleproduction. Third, there will be fluctuations created in theUniverse in the wake of this transition. The first featurehas been examined in great detail in the cosmologyliterature. For the last two features, there are many quantumfield theory calculations of fluctuation and/or dissipationdynamics [11–15,15–21] but very little has been explicitlyapplied to particle physics models during cosmologicalphase transitions. One exception is in the case of inflation,where warm inflation captures all three of these features[22–33]. However, cosmological phase transitions can andgenerically do occur with no inflation and, in these casesalso, all three of these features will be present. They are anintrinsic part of the evolution history of the early Universeand the dynamics emerging from whatever is the under-lying particle physics model.Fluctuation-dissipation effects will, more generally, be

present in the dynamics of any cosmological scalar field,regardless of the occurrence of phase transitions. Forexample, several completions of the SM predict theexistence of very light scalars, such as extradimensionalmoduli or axionlike fields. These fields will be

underdamped during the early inflationary phase and drivento potentially very large values by random quantumfluctuations. After inflation, once the expansion rate hasdecreased sufficiently, they will be able to dynamicallyrelax to their minimum energy configuration. This may inseveral cases lead to large-amplitude oscillations thatoverclose the Universe or spoil the successful predictionsof big bang nucleosynthesis (BBN) for light elementabundances, which poses a considerable challenge forcosmological models in beyond the SM scenarios.Interactions with other degrees of freedom in the ambientheat bath may, however, induce energy dissipation andfluctuations in these scalar fields, modifying their dynami-cal evolution and potentially their role in the subsequentcosmic history.Another case where fluctuation-dissipation dynamics

may be of relevance is the cosmological variation offundamental constants driven by scalar fields. These couldinclude e.g. unknown scalars driving variations of the fine-structure constant, α, or even the cosmological evolution ofthe SM Higgs field, which determines all fermion massesand in particular the electron-proton mass ratiome=mp (seee.g. [34]). In the latter case dissipation could delay theelectroweak phase transition, as we will discuss in thiswork for generic phase transitions, and potentially yieldtemporal variations of the electron-proton mass ratiome=mp. Additionally, the associated fluctuations will alsoinduce spatial variations of this ratio, which depending ontheir size and scale could in principle lead to observableeffects.This fluctuation-dissipation dynamics is not specific to

near thermal equilibrium conditions. Whatever the statis-tical state is, a relation exists between the dissipationproduced by the system and the fluctuations induced bythe radiation bath. The near thermal equilibrium regime is,however, amenable to explicit calculations using well-developed thermal field theory methods and will be thefocus of this paper. The early Universe is generally believedto be in a near thermal equilibrium state and so thesecalculations based on thermal field theory have significantrelevance to it. Nevertheless, there could be processes in theearly cosmic stages where a scalar field moves too quicklyor the underlying microphysical processes are too slow tojustify a near thermal equilibrium approximation. Thus itshould be kept in mind that the calculations done in thispaper could also be extended to these regimes; it would be atechnical, albeit complicated, step further, but the under-lying concept is the same as that developed in this paper.It is the goal of this work to set the stage for the study of

cosmological fluctuation-dissipation dynamics within con-crete particle physics models. We discuss different exam-ples of dissipation (and related noise) coefficients withinthe SM and beyond, exploring their distinct parametricregimes and domains of applicability. We then outlinesome of the generic consequences of dissipation, particle

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production and induced fluctuations in the dynamics ofcosmological scalar fields, both with and without theoccurrence of phase transitions. We focus on the postinfla-tionary dynamics, where the effects of fluctuation-dissipation dynamics remain largely unexplored, thereexisting already a considerable literature devoted to thistopic in inflationary cosmology in the context of the above-mentioned warm inflation dynamics.To better illustrate the cosmological impact of these

processes, we analyze in detail two concrete scenarios.First, we consider a high-temperature phase transition inthe early Universe, where the associated Higgs field candissipate its energy into fermionic modes through standardYukawa interactions. In particular, we show that, byslowing down the field’s motion, dissipation will delaythe phase transition, leading to additional entropy produc-tion and Hubble expansion that can dilute the abundance ofdangerous relics such as topological defects. Furthermore,if the transition is sufficiently delayed, the Higgs field maycome to dominate the energy balance and yield an addi-tional (short) period of inflationary expansion. This resultsin a more efficient dilution of thermal relics, similarly tothermal inflation models, although dissipative friction cansustain accelerated expansion below the temperatures atwhich thermal effects can hold the Higgs field in thesymmetric phase. Both thermal and dissipative (warm)inflation are, in fact, due to the same interactions betweenthe Higgs field and the thermal bath degrees of freedom andmay occur within the same cosmological phase transition,as we show in this work.Second, we consider the relaxation of a scalar field from

a large postinflationary value to its minimum energyconfiguration. We show that its coupling to a B- or L-violating sector can lead to the dissipative production of abaryon or lepton asymmetry, respectively, in the spirit ofthe warm baryogenesis scenario proposed in [35] in theinflationary context. To illustrate this generic mechanism,we develop a concrete model of dissipative leptogenesis,where dissipation results from the excitation and decay ofheavy right-handed neutrinos, which gain a large Majoranamass from the coupling to a dynamical scalar field. Asopposed to standard leptogenesis and other thermal baryo-genesis scenarios, these are mainly produced off shell,which allows for baryogenesis at parametrically lowtemperatures and therefore avoids the troublesome over-production of thermal relics.These two examples show that dissipative effects can

have an important role in the cosmic history, particularly inaddressing some of the most important puzzles in moderncosmology. We therefore hope that they motivate a morethorough exploration of this topic and of more generalnonequilibrium processes in cosmology.This work is organized as follows. In Sec. II we show

examples of how dissipation arises within common particlephysics models, focusing in particular on the electroweak

phase transition and grand unified theories. In Sec. III wedescribe the effects of fluctuation-dissipation dynamics inhigh-temperature phase transitions. In Sec. IV we describethe postinflationary production of a baryon asymmetrythrough dissipative effects, describing in detail the dis-sipative leptogenesis scenario and its observational signa-tures. We summarize our main results and conclusions inSec. V, also discussing prospects for future research inthis area.

II. FLUCTUATION-DISSIPATION DYNAMICSIN PARTICLE PHYSICS MODELS

The several scalar fields employed in particle physicsand associated cosmological models are typically notisolated systems and generically interact with other degreesof freedom. Their dynamics is therefore described by aquantum effective action that encodes the effects of allinteractions with other fields. In the cosmological context,this effective action must take into account the nontrivialstatistical states of both the dynamical field and the degreesof freedom with which it interacts. The black-body spec-trum of the cosmic microwave background (CMB) andthe successful predictions of BBN show that the Universewas in a state very close to local thermal equilibrium fora great part of its early history, and we henceforth assumethat all relevant particle states always remain near thisconfiguration.For static fields the effective action reduces to a local

effective potential, which takes the well-known Coleman-Weinberg form at leading order in a perturbative expansion[36]. From the finite temperature effective potential one canderive the thermodynamic properties of the cosmologicalfields, such as their energy density, entropy and pressure, aswell as thermal mass corrections. Static fields are, however,generically of little interest in cosmology, and for dynami-cal fields the effective action includes nonlocal effectsbeyond the leading effective potential approximation.Time nonlocal effects may take different forms depend-

ing on the regime considered. The simplest case is theadiabatic regime, where the field varies on time scales thatlargely exceed the typical time scales of the relevantmicrophysical processes. This is for example the casefor the inflaton field, which in the simplest scenarios isslowly rolling in order to produce a quasi–de Sitter phase.Local thermal equilibrium in the ambient heat bath can bemaintained if scattering and/or decay processes within it aresufficiently fast, namely faster than Hubble expansion, suchthat an adiabatic approximation will be valid for largeclasses of cosmological scalar fields.In the adiabatic limit, a system has sufficient time to

relax to an equilibrium configuration in response to theperturbing effect of the time nonlocal terms in the effectiveaction, and linear response theory can be used to study thesystem’s evolution. The leading time nonlocal effect isdissipation of the scalar field’s energy into the degrees of

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freedom in the heat bath, which manifests itself through aneffective friction term in the field’s equation of motion.The simplest example of this is the creation and

subsequent annihilation of particle-antiparticle pairscoupled to the background field, where this couplingmakes the amplitude of creation and annihilation fielddependent. Suppose then that pairs are created at a time twhere the scalar field takes a value ϕðtÞ. They will thenannihilate at time tþ δt where the field has shifted by anamount δϕ ¼ _ϕδtþ… in the adiabatic regime, and toleading order there will be a net particle productionproportional to _ϕ, resulting in a transfer of energy fromthe scalar field into the produced degrees of freedom. Thiswill perturb the local thermal equilibrium in the ambientheat bath but the system can relax into a new equilibriumconfiguration if adiabatic dissipation is slower than othermicrophysical processes.Dissipation corresponds to the systematic effect of the

particles in the heat bath on the evolution of the field andthe resulting friction opposes the latter’s evolution throughthe creation and annihilation of particles in the heat bathin a field-dependent fashion, as outlined above. This isentirely analogous, for example, to the systematic frictionforce produced on a moving mirror by a rarefied gas ofmolecules that randomly hits the mirror in a Brownianmotion [37]. Much like this random Brownian motion also

resulting in an irregular motion of the mirror, fluctuationsin the cosmological heat bath will also backreact on theevolution of the scalar field and introduce a degree ofrandomness. The two effects, fluctuations and dissipation,result from the same interactions between the scalar fieldand the heat bath and are thus interconnected. This is ageneral result that applies to large classes of dissipativesystems in nature and is known as the fluctuation-dissipation theorem, the details of which depend on thestatistical state of the system and its microscopic properties.The combined effect of fluctuations and dissipation leads

to an effective Langevin-like equation for a cosmologicalscalar field interacting with the ambient heat bath of theform (see e.g. [15,21,27,38–40] and references therein)

ϕþ ð3H þϒÞ _ϕ −1

a2∇2ϕþ V 0ðϕÞ ¼ ζ; ð1Þ

where ϒ denotes the dissipation coefficient and ζ therelated random noise term, with the remaining termsyielding the usual Klein-Gordon equation in a flatFriedmann-Robertson-Walker universe. In the adiabaticregime, the noise term is Gaussian to leading order andits correlator satisfies the following fluctuation-dissipationrelation in momentum space [33,41,42]:

hζðk; tÞζðk0; t0Þi ¼�ð1þ 2nðkÞÞ 3H

2ffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi1þ 4πQ=3

pπ

þ 2ϒT

�a−3ð2πÞ3δ3ðkþ k0Þδðt − t0Þ; ð2Þ

where we have also included a “quantum noise” contribu-tion, given by the first term within brackets andQ ¼ ϒ=3H. This results from a coarse-graining of thescalar field as employed in the stochastic approach toinflation [43], with short-wavelength (subhorizon) fieldmodes inducing an effective noise in the dynamics of thelong-wavelength classical modes. While the form in Eq. (2)is obtained for a sharp mode splitting, a smooth filteringfunction generically results in a colored noise distribution[44]. We have also included the effect of a generic phase-space mode distribution nðkÞ [42], which vanishes in thestandard stochastic inflation approach, but becomes sig-nificant, in particular, when the scalar field is itselfthermalized and nðkÞ is the Bose-Einstein distribution.For example, for T ≫ H; jV 00ðϕÞj andQ ≪ 1, the first termwithin brackets becomes proportional to HT for modescrossing the horizon.As discussed above, the dissipative friction term is

associated with a net particle creation in the ambient heatbath. One can then integrate Eq. (1) and average over thenoise term to obtain the evolution of the scalar field’senergy density, and use energy conservation to derive theassociated equation for the heat bath:

_ρϕ þ 3H _ϕ2 ¼ −ϒ _ϕ2; _ρα þ 3Hðρα þ pαÞ ¼ ϒ _ϕ2;

ð3Þwhere we take the heat bath to be described, to leadingorder, by a perfect fluid of density ρα and pressure pα. If itis composed of relativistic particles that thermalize suffi-ciently fast, the latter corresponds to a radiation fluid withequation of state pR ¼ ρR=3.As we discuss below in more detail, we will be mainly

interested in interactions between cosmological scalarfields and other (complex) scalar and fermionic degreesof freedom. Gauge interactions may also be of relevancefor early universe cosmology, but since the main featuresof vector boson interactions are well described by scalardegrees of freedom we will not consider this case explicitlyto simplify our discussion. We thus consider a generic(renormalizable) Lagrangian of the form

L ¼ fðϕÞjχj2 þ gϕψψ ; ð4Þwhere χ and ψ denote complex scalar and fermion fields inthe heat bath and fðϕÞ is a generic function of thedynamical scalar field we are interested in.

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The leading one-loop contributions of these interactionsto the effective action are illustrated in Fig. 1 and, for anearly thermal heat bath at temperature T ¼ 1=β, theseyield an adiabatic dissipation coefficient of the form

ϒ ¼ f0ðϕÞ2T

Zd4pð2πÞ4 ρχðp0;pÞ2nBðp0Þð1þ nBðp0ÞÞ

þ g2

2T

Zd4pð2πÞ4 Tr½ρψðp0;pÞ2�nFðp0Þð1 − nFðp0ÞÞ;

ð5Þ

where nBðωÞ ¼ ðeβω − 1Þ−1 is the Bose-Einstein andnFðωÞ ¼ ðeβω þ 1Þ−1 is the Fermi-Dirac distribution forparticle modes of energy ω. The functions ρχ and ρψrepresent the spectral functions of the scalar and fermionfields in the heat bath and can be computed from thecorresponding (dressed) propagators at finite temperature,using e.g. the real-time formalism [33,45–47]. For exam-ple, in the scalar field case one obtains

ρχ ¼4ωpΓχ

ðp20 − ω2

pÞ2 þ 4ω2pΓ2

χ; ð6Þ

where ωp ¼ffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffijpj2 þm2

χ

q, with the field mass correspond-

ing to its renormalized value including thermal corrections,and Γχ is the (finite temperature) decay width of the field. Asimilar, albeit more complicated expression can be obtainedin the fermionic case [46].From Eq. (5) one can deduce a few generic aspects of

dissipation coefficients in the adiabatic regime. First, wesee that if the fields χ and ψ were in a trivial (vacuum) state,the dissipation coefficient would vanish. Dissipation is thusan effect intrinsic to the motion of the scalar field in thepresence of a heat bath with nontrivial occupation numbers,corresponding as argued above to the systematic effect ofthe heat bath degrees of freedom on the field’s motion.Second, the spectral functions correspond to the imaginarypart of the field propagators and are consequently propor-tional to their decay width, as shown above. Hence, if thefield’s χ and ψ were stable there would be no dissipation.

However, at finite temperature interacting fields alwayshave a nonzero decay width, arising from a combination ofdecays, inverse decays and Landau damping processes.Finally, the dissipation coefficient will in general be bothfield and temperature dependent. The temperature depend-ence is explicit in the distribution functions but will alsoarise in general in the masses and decay width of the fields.The field dependence can be explicit in the scalar case, for ageneric function fðϕÞ, but will also arise from the masses(and consequently the decay width of the fields), notingthat from Eq. (4) one obtains at tree level m2

χ ¼ fðϕÞand m2

ψ ¼ g2ϕ2.The integrals in Eq. (5) can be computed numerically in

general, there existing, however, two approximate regimeswhere the computation can be performed analytically. In thelow-temperature regime,mχ ; mψχ

≫ T, the distribution func-tions nB; nF become exponentially (Boltzmann) suppressedfor on-shell field modes, p2

0 ¼ ω2p, so that their contribution

can be neglected. The main contribution in this case comesfrom off-shell or virtual modes with jpj; p0 ≪ mχ ; mψ , forwhich the spectral functions take a simple form, e.g. ρχ ≃4Γχ=m3

χ in the scalar case. The low-momentum dissipationcoefficient (LM) is thus approximately given by

ϒLM ≃ 16f0ðϕÞ2T

Zd4pð2πÞ4

Γ2χ

m6χnBð1þ nBÞ

þ 2g2

T

Zd4pð2πÞ4

TrðImΣ2Þm2

ψnFð1 − nFÞ ð7Þ

where ImΣ is the imaginary part of the fermion self energy,from which their decay width can be extracted in theconventional way. This approximation is valid in the narrowwidth limit where mi ≫ Γi for i ¼ χ;ψ (see [47] for moredetails). The integrals involving the distribution functions andthe decay widths can then be performed numerically [46,47].A simple example that we will consider below is the casewherefðϕÞ ∝ ϕ2,withmχ ∝ ϕ andΓχ ∝ mχ . In this case, it isnot difficult to see that ϒLM ∝ T3=ϕ2 for the scalar contri-bution, while the corresponding fermionic contribution issuppressed by further powers ofT=mψ ≪ 1 as shown in [46].In the opposite high-temperature regime, mχ ; mψ ≪ T, it

is energetically possible to excite on-shell modes in thethermal bath and their occupation numbers are notBoltzmann suppressed. These will then give the dominantcontribution to the dissipation coefficient, and one canexpand the spectral functions about their poles at p0 ¼ ωp

to yield

ϒP ≃ f0ðϕÞ2T

Zd3pð2πÞ3

1

Γχω2pnBð1þ nBÞ

þ 2g2

T

Zd3pð2πÞ3

m2ψ

Γψω2pnFð1 − nFÞ: ð8Þ

FIG. 1. Leading one-loop contributions of scalar and fermiondegrees of freedom to the dissipation coefficient. The thickdashed and solid lines indicate dressed propagators for scalarsand fermions, respectively.

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The 3-momentum integrals can then be easily computedanalytically in different regimes (see e.g. [47]). In particu-lar, for light on-shell modes one typically obtainsΓi ∝ mi ∝ T, for i ¼ χ;ψ, yielding ϒP ∝ ϕ2=T for scalarmodes and ϒP ∝ T for fermionic modes [46,47].In the general case, the dissipation coefficient receives

contributions from both on-shell and off-shell modes, andnumerical calculations show that adding both contributionsyields a very good approximation to the full result. Inparticular, it has been observed that the on-shell contribu-tion can be dominant for masses mi ≳ T despite theassociated Boltzmann suppression, particularly for smalldecay widths [47].These two regimes will be relevant for different types of

particle physics and cosmological scenarios. On the onehand, in a typical phase transition the relevant Higgs field isstabilized at the origin at high temperatures and startsrolling towards the minimum of its potential below a criticaltemperature (potentially after tunneling in a first orderphase transition). In this case, the fields it couples to areinitially light, and on-shell dissipation dominates. As Tdecreases and the field value approaches the true minimum,these fields become heavier and the contribution of low-momentum modes will grow until it potentially dominates.On the other hand, a light scalar ϕ can attain very largevalues during inflation, after which it will eventually rolltowards the minimum of its potential. In this case, off-shellmodes will typically dominate initially, while on-shellmodes will become increasingly more significant if ϕevolves towards smaller values and χ;ψ become lighter.In the following we give a series of examples, by no

means exhaustive, of dissipation coefficients for dynamicalscalar fields in the SM and its typically consideredextensions, within different dynamical regimes that maybe relevant for the cosmic history.

A. Dissipation in the SM andsupersymmetric extensions

The SM gauge group, SUð3Þc × SUð2ÞL × Uð1ÞY , isbroken spontaneously to SUð3Þc ×Uð1ÞQ by the non-vanishing vacuum expectation value of the electroweakHiggs boson. In typical cosmological scenarios, the reheat-ing temperature after inflation largely exceeds the criticaltemperature of the electroweak phase transition. Quarks,leptons and electroweak gauge bosons are relativistic andin thermal equilibrium, and their backreaction on theHiggs effective potential at high temperatures stabilizesthe Higgs field at the symmetric minimum. As the Universecools down, the effective potential approaches its zero-temperature form and the Higgs field will roll towards thefinite vev that spontaneously breaks the electroweaksymmetry. As the field rolls from the origin towards thebroken minimum, we then expect dissipative processes tobe mainly mediated by on-shell quarks and leptons, as wellas the weak gauge bosons. The former, in particular, havethe following well-known Yukawa couplings to the Higgsfield:

L ∼ λije eR;iϕ†Lj þ λiju uR;iϕqj þ λijd dR;iϕ†qj þ H:c:; ð9Þ

where we have suppressed weak isospin and color indices,while the indices i; j label the fermion generations. At hightemperatures, the decay width of quarks and leptons isgiven essentially by Landau damping terms from the aboveYukawa interactions, as well as gauge interactions [48].The relevant dissipation coefficient is thus of the on-shellform given in Eq. (8). Including both Yukawa and gaugecontributions to the fermions’ decay width we obtain

ϒPSM ≃ 288ζð3ÞT

π3X3i¼1

� ðλiie Þ2ðλiie Þ2 þ 4ðg21 þ 3g22Þ

þ 3ðλiiu Þ2ðλiiu Þ2 þ 4ðg21 þ 3g22 þ 8g23Þ

þ 3ðλiid Þ2ðλiid Þ2 þ 4ðg21 þ 3g22 þ 8g23Þ

�; ð10Þ

where gi are the SM gauge couplings. We note thatdissipation can also occur through the excitation of theW� and Z gauge fields, although for simplicity we do notinclude this in the above expression. As ϕ increases and Tdecreases eventually the masses of the SM particles willbecome heavier than the temperature and their on-shellcontribution to dissipation becomes Boltzmann suppressed,with low-momentum dissipation of the form in Eq. (7)becoming the dominant contribution.Another example based on the same symmetry group

arises within the minimal supersymmetric extension of theSM (MSSM), where two Higgs doublets are required tobreak the electroweak symmetry and give masses to allquarks and leptons. The MSSM superpotential is given by

W ¼ μHuHd þ yuHuQUc þ ydHdQDc þ yeHdLEc:

ð11Þ

As in the SM, in this case one finds dissipative channels forboth Higgs scalar components, hu and hd, by exciting bothfermion and sfermion degrees of freedom in the heat bath.At high temperatures these will have the forms ϒ ∝ ϕ2=Tand ϒ ∝ T obtained above for sfermions and fermions,respectively, with ϕ ¼ hu; hd (see also [49]).The simplest extension of the MSSM, known as the next-

to-minimal supersymmetric SM (NMSSM), replaces the μterm in the superpotential by a trilinear term gΦHuHd,where Φ is a singlet chiral superfield. The effective μ term

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is then given by the vev of the scalar component of Φ, apossibility that helps address the smallness of the a prioriunconstrained μ parameter required for successful electro-weak symmetry breaking (see e.g. [50]). One can thenenvisage scenarios where the singlet scalar field, ϕ, isdriven towards (or maintained at) a large value duringinflation, after which it will roll towards hϕi ¼ μ=g. Itscoupling to both Higgs scalar doublets hu and hd is of theform given in Eq. (4) and explicitly as

LS ¼ g2jϕj2ðjhdj2 þ jhuj2Þ þ gϕ†h†dyiju ~qi ~ucj

þ gϕ†h†uðyijd ~qi ~dcj þ yije ~li ~ecjÞ þ H:c:; ð12Þ

so that the singlet field can dissipate its energy throughexcitation of both scalar doublet components and theirfermionic superpartners, which decay into the SM fermionsand sfermions. If the initial field value is large, thedominant contribution to the dissipation coefficient is givenby off-shell scalar modes as discussed above and thedissipation coefficient is approximately given by

ϒLMNMSSM ≃ Cϕ

T3

ϕ2;

Cϕ ≃ 1

8π

Xij

ð3ðyiju Þ2 þ 3ðyijd Þ2 þ ðyije Þ2Þ; ð13Þ

where i; j run over family indices.

B. Dissipation in grand unified theories:an SUð5Þ example

There is significant evidence for the unification of theSM gauge couplings at high-energy scales, particularlywithin the context of the MSSM [51–54], which pointstowards the existence of a larger gauge symmetry group.Several GUT have been proposed where this gauge

group is spontaneously broken into the SM gauge groupthrough a Higgs-like mechanism, with SUð5Þ and SOð10Þbeing the simplest and most studied examples (see e.g.[55]). In GUT models the relevant Higgs fields are coupledto gauge bosons and matter fields, such that fluctuation-dissipation dynamics may play an important role in theircosmological evolution.If GUT symmetries are restored after inflation, for a

sufficiently high reheating temperature, the Higgs fieldsroll from the symmetric point to the symmetry-breakingminimum once the temperature drops below a criticalvalue. This may be preceded by a tunneling event if thetransition is first order, depending on the particle content ofthe GUT model [2], but dissipative rolling will alwaysoccur. This scenario may, however, be troublesome forcosmology since symmetry breaking typically leads tothe generation of dangerous topological defects such as

monopoles, which may overclose the Universe. Althoughpostinflationary symmetry restoration is appealing from thepoint of view of thermal GUT baryogenesis models, thereare viable alternative mechanisms for the production of acosmological baryon asymmetry such as the dissipativebaryo/leptogenesis mechanism that we describe in Sec. IV.In the case where GUT symmetries are not restored

during reheating, the relevant Higgs fields may neverthelessfind themselves displaced from the symmetry-breakingminimum after inflation. This occurs if the Higgs fieldsare light during inflation, being frozen at some initial valueor even driven to larger values by random quantumfluctuations. Dissipation will then also be relevant in thepostinflationary eras as the fields roll towards the trueminimum of their potential.To illustrate the form of dissipative effects in GUT

models, we consider the simplest case of SUð5Þ, bearing inmind that similar processes will generically occur forhigher-rank gauge groups where many other dynamicalscalars and dissipative channels may be present. In fact, inSec. IV we consider the particular example of a scalar fieldresponsible for the Majorana mass of right-handed neu-trinos that is naturally embedded in SOð10Þ models.SUð5Þ is broken into the SM gauge group by the vev of

an adjoint Higgs field, 24H, which gives masses to thegauge and fundamental Higgs field components that areassociated with the broken symmetries. The adjoint scalarpotential takes the form

Vð24HÞ ¼ −μ2Tr½242H� þ aTr½242H�2þ bTr½244H� þ cTr½243H�; ð14Þ

which, in certain parametric regimes, has an absoluteminimum in the direction ϕdiagð2; 2; 2;−3;−3Þ= ffiffiffiffiffi

30p

thatpreserves the SM gauge group. Interactions between theadjoint and fundamental Higgs fields are given by

Ls ¼ −A2

25†H5H þ B

4ð5†H5HÞ2 þ C5†H5HTr½242H�

þD5†H242H5H þ E5†H24H5H; ð15Þ

while the latter is coupled to the SM matter fermions in the10 and 5 representations via Yukawa couplings of the form

LY ¼ Yij5 5Fi10Fj5

�H þ 1

8ϵ5Y

ij1010Fi10Fj5H þ H:c: ð16Þ

Decomposing these fields in terms of SM representations,we find the following interaction Lagrangian involving thesymmetry-breaking scalar direction ϕ, the doublet andtriplet Higgs fields, H and T, and the SM quarks andleptons:

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Lint ¼�−A2

2þ 2Effiffiffiffiffi

30p ϕþ 2D

15ϕ2 þ Cϕ2

�jTj2 þ

�−A2

2−

3Effiffiffiffiffi30

p ϕþ 3D10

ϕ2 þ Cϕ2

�jHj2

− Yij5 ðliqj þ dci u

cjÞT� − Yij

10

�1

2qiqj þ uci e

cj

�T þ Yij

5 ðLiecj þ qidcjÞH� − Yij10qiu

cjH þ H:c: ð17Þ

As discussed above there are also gauge interactions, butfor illustrative purposes we restrict ourselves to dissipativeeffects associated with the scalar and Yukawa interactionsgiven above. Let us consider, in particular, the case where ϕhas a large vev after inflation that is displaced from its trueminimum. The leading C terms in Eq. (17) give a largemass to the Higgs doublet and triplet, which are initiallyequal due to the large field vev, while a doublet-triplet masssplitting will only arise close to the minimum. These termsare of the generic form given in Eq. (4) for χ ¼ H; T and weexpect the main contribution to dissipation in this regime tocorrespond to virtual Higgs modes decaying into quarksand leptons. From Eq. (7) we then obtain the followingdissipation coefficient:

ϒLMSUð5Þ ≃

0.44C2

T7

ϕ6

Xi;j

½10ðYij5 Þ4 þ 8ðYij

10Þ4�: ð18Þ

In a supersymmetric realization of SUð5Þ the SM fermionshave scalar superpartners and, due to the holomorphicnature of the superpotential, two distinct Higgs fields in the5 and 5 representation are required. The relevant part of thesuperpotential is given by

W ¼ g5H24H5H þM5H5H þ Yij5 5Fi10Fj5H

þ 1

8ϵ5Y

ij1010Fi10Fj5H; ð19Þ

where the relevant scalar interactions are

Ls ¼ Yij5

�2gϕffiffiffiffiffi30

p þM

�~tu†ð ~dci ~ucj − ~li ~qjÞ

þ Yij10

�2gϕffiffiffiffiffi30

p þM

�~td†�~eci ~u

cj þ

1

2~qi ~qj

�

þ Yij10

�M −

3gϕffiffiffiffiffi30

p�~h†d ~qi ~u

cj

þ Yij5

�M −

3gϕffiffiffiffiffi30

p�~h†uð~li ~ecj þ ~qi ~d

cjÞ þ H:c: ð20Þ

As in the nonsupersymmetry (SUSY) model, the low-temperature regime for dissipation will be the relevant oneafter inflation if ϕ attains a large vev, gϕ ≫ M, and theGUT symmetry is not restored. Dissipation is in this casedominantly mediated by virtual scalar doublet and tripletHiggs modes that decay mainly into sfermion fields, as

shown in [46] for generic SUSY models of this form. Thedissipation coefficient is then given by

ϒLMSSUð5Þ ≃

1

16π

T3

ϕ2

Xi;j

½10ðYij5 Þ2 þ 8ðYij

10Þ2�: ð21Þ

Note that dissipative effects will be more pronouncedin this case compared to the non-SUSY model, since thedissipation coefficient is less suppressed by powers ofT=mχ , where χ generically denotes the doublet and tripletHiggs scalars involved.

III. FLUCTUATION-DISSIPATION DYNAMICS INCOSMOLOGICAL PHASE TRANSITIONS

Despite the numerous studies in the context of con-densed matter systems, the dynamics of phase transitionsin fundamental particle physics and cosmology remainslargely unexplored. The recent discovery of the Higgsboson at the LHC, so far consistent with the SM predictionsfor the spontaneous breaking of the electroweak gaugesymmetry, is the first experimental hint for the occurrenceof a fundamental phase transition in the cosmic history andmotivates further exploration of this topic. Moreover, theapparent unification of gauge couplings suggests, as dis-cussed above, that one or more phase transitions may haveoccurred in the early stages of the Universe’s history,spontaneously breaking a higher-ranked gauge group intoSUð3Þc × SUð2ÞL ×Uð1ÞY , in potentially several stages ofa progressively lower degree of symmetry.Cosmological Higgs fields are coupled to matter fields

and gauge bosons, and the effects of dissipation andassociated fluctuations will necessarily play a role in theevolution of the fields from a symmetric to a spontaneouslybroken symmetry phase. In this section, we discuss severalpotential effects of fluctuation-dissipation dynamics ingeneric cosmological phase transitions.

A. Thermal fluctuations and topological defects

The cosmological evolution of a generic Higgs field inthe process of spontaneous symmetry breaking follows aLangevin-like equation of the form (1), with both the noiseterm on the right-handed side and the dissipative frictionterm on the left-handed side playing an important role atdifferent dynamical stages. Fluctuations will be primarilysignificant at the onset of the phase transition, just belowthe critical temperature at which the symmetric Higgs valuecan no longer be stabilized by thermal effects. In particular,

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in the absence of random fluctuations the field wouldremain at the unstable symmetric minimum, since this isnevertheless an extremum of the effective potential. Thenoise term in the Langevin-like equation is thus crucial ininducing the phase transition and in determining thedirection within the vacuum manifold towards which thefield’s evolution will proceed.Since the Higgs field is, on average, at rest at the onset of

the phase transition, its dynamics will be initially governedby the Gaussian and white noise term in the adiabaticregime. As discussed earlier, the stochastic noise termencodes the effective backreaction of the ambient heat bath,also incorporating the inherent quantum nature of the field.On the one hand, the backreaction of the heat bath isdirectly related to the dissipation coefficient through thefluctuation-dissipation theorem and given by the termproportional to ϒ in the noise correlator (2). On the otherhand, the remaining quantum noise term, proportional tothe Hubble parameter H, can be deduced from a coarse-graining of the quantum Higgs field, with short wavelengthmodes that are well within the Hubble horizon backreactingon the longer wavelength modes that one is interested infollowing. This stochastic approach has e.g. been success-ful in describing field fluctuations in both warm and coldinflation regimes [41].The phase transition will then initially be driven by the

quantum and thermal/dissipative noise terms, which ran-domly kick the Higgs field away from and towards thesymmetric minimum. This will proceed until the amplitudeof the noise term becomes subdominant compared to theclassical terms in the equation of motion, i.e. roughly whenffiffiffiffiffiffiffiϒT

p ≲ V 0ðϕÞ=H2 for strong dissipation. Random fluctua-tions will then effectively cease and the subsequent fielddynamics will essentially be classical. However, the field isnow spatially inhomogeneous and the classical evolutionwill drive it to different directions in the vacuum manifoldat distinct spacetime points. The classical dynamics cannevertheless homogenize the field within causally con-nected patches, determined by the field’s correlation length,ξc, that is at most the size of the cosmological horizon. Ifthe Higgs field is relativistic at this stage, ξc ∼ 1=T ≪ H−1,and so the temperature at which the noise term becomesinefficient will set the size and consequently the abundanceof any topological defects that may form once the fieldsettles into the lowest energy configuration. Some prelimi-nary studies for the formation of topological defects inphase transitions including the effects of both thermal noiseand dissipation have been performed in [56–58]. It wouldbe interesting to further explore this in the context of theconcrete particle physics models discussed above andwithin realistic cosmological settings. This is, however,beyond the scope of the present work, where we focus onthe dissipative classical evolution.One other related consequence of the noise term should

nevertheless be pointed out. The inhomogeneity of the

Higgs field resulting from its initial random motion willalso induce a spatial variation of its gauge quantumnumbers, generically sourcing magnetic fields [5,59].Their strength will then also be determined by the corre-lation length at the time when the noise term becomesinefficient and, if sufficiently large, this may sow the seedsfor galactic magnetic fields.Although here we will not pursue these issues in further

detail, it is worth emphasizing that the Langevin-likeequation (1) gives a fundamental framework for thesestudies. Given a particle physics model, one can computethe dissipation coefficient and associated noise term fromfirst principles, as explicitly done in the previous section forseveral examples, and use this equation to determine boththe quantum and classical dynamics. This allows one todetermine the correlation length, the density of cosmologi-cal defects or the strength of generated magnetic fieldsin a rigorous way. In this way, there is no need to simplyemploy statistical arguments to derive these quantities andthe field evolution can be completely determined forarbitrary initial conditions.

B. Dissipative effects: entropy production andadditional inflation

Once the effects of the thermal and quantum noisesbecome subdominant, the field’s evolution becomesclassical and is driven by the competition between thescalar potential’s slope and the effects of dissipative andHubble friction. To analyze the concrete effects ofdissipation, which have so far been overlooked in theliterature and, as we will show, may play an importantrole, we consider a generic toy model where a real Higgsfield is coupled to fermions through standard Yukawacouplings of the form in Eq. (4). This can be easilyextended to concrete particle physics models such as theelectroweak phase transition or GUT phase transitions byconsidering the appropriate couplings, particle contentand the properties of the vacuum manifold, as illustratedin the previous section.The fermions induce, as discussed before, both local

and nonlocal corrections to the effective action of theHiggs field. The leading effect of the former is finitetemperature corrections to the effective potential, withzero-temperature corrections playing a subdominant rolethat we will for simplicity discard in our analysis.Thermal corrections are significant for relativistic fer-mions, mψ ≪ T, namely inducing a thermal mass for theHiggs field, while for mψ ≳ T these corrections areBoltzmann suppressed and thus irrelevant to the dynam-ics. The general form of the thermal mass can be obtainedby numerical integration, but for our purposes it issufficiently accurate to explicitly multiply the high-temperature result by a Boltzmann factor, yielding forthe effective Higgs potential

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Vðϕ; TÞ ¼ λ2

4ðϕ2 − v2Þ2 þ 1

2α2T2ϕ2 exp

�−mψ ;T

T

�;

ð22Þ

where α2 ¼ g2NF=6, with NF denoting the number ofDirac fermion species. In this expression, the first termcorresponds to the simplest symmetry-breaking potentialwith minima at ϕ ¼ �v and the second term is the leadingthermal correction. In the fermion mass m2

ψ ;T ¼ g2ϕ2 þh2T2 we also include thermal corrections from couplingto different species in the heat bath, including e.g. gaugefields, which we generically parametrize with an effectivecoupling h≲ 1. At high temperatures, the second termdominates the effective potential and the Higgs field isthus stabilized at the origin, with mψ ;T ≃ hT ≲ T, while atlow temperatures it will roll towards one of the minima atϕ ¼ �v. The symmetric minimum becomes unstable at acritical temperature:

Tc ¼λvα: ð23Þ

As soon as the field begins rolling towards the symmetry-breaking minimum, it will feel the friction effect of thefermion heat bath. The relevant dissipation coefficientcorresponds to the on-shell excitation of light fermionswith a thermal decay width and was first computed in[60], yielding

ϒ≃ 11.2NFT exp

�−mψ ;T

T

�: ð24Þ

Note that, as above, we have multiplied the high-temperature result by a Boltzmann factor which willcut off on-shell dissipation at low temperatures, T ≲ gϕ,which is sufficiently accurate for our purposes. As wehave seen above, virtual modes will also induce dissi-pation in the latter regime, but since this is a significantlysmaller effect we discard it in our analysis to a firstapproximation. Also notice that the SM high-temperaturedissipation coefficient given in Eq. (10) coincides withthis expression if one discards gauge interactions. Thethermal width of the fermions is roughly given byΓψ ∼m2

ψ ;T=T ≲ T, and adiabaticity of the dissipativeprocess requires Γψ ≳H, which is easily satisfied in

the radiation-dominated era, where H ≃ ffiffiffiffiffiffiffiffiffiffiffiffiπ2=90

pg1=2� T2=

MP ≪ T. For simplicity, we assume that the radiationbath is made exclusively of fermions, taking g� ¼ 7NF=4,although one can easily include other relativistic degreesof freedom.With the form of the effective potential and dissipation

coefficient, we may thus describe the dynamics of the phasetransition by solving the system of coupled Higgs-radiationequations, given by

ϕþ ð3H þϒÞ _ϕþ V 0ðϕÞ≃ 0; _ρR þ 4HρR ¼ ϒ _ϕ2;

ð25Þwhere as discussed abovewe neglect the effects of the noiseterm, and ρR ¼ ðπ2=30Þg�T4 assuming a nearly thermalequilibrium state. Dissipation thus plays two distinct rolesin the dynamics, on the one hand damping the field’smotion and, on the other hand, sourcing the radiation baththrough the production of fermion modes.If the energy density in the scalar field is sufficiently

large, it may come to dominate the energy density beforethe Universe cools down to below the critical temperature.This occurs for g2NF ≳ 2π

ffiffiffiffiffig�

pλ, thus inducing a period of

thermal inflation along the lines proposed in [61] and whichmay help diluting the abundance of dangerous thermalrelics produced e.g. during reheating or earlier cosmologi-cal phase transitions. This additional period of inflationcan typically last only for a few e-folds until the criticaltemperature is reached, with the field then rolling towardsthe symmetry-breaking minimum and oscillating about it.In the presence of dissipation, the dynamics can be quite

different below the critical temperature and an interestingalternative/addition to thermal inflation arises. First, wenote that in the radiation era ϒ≳ T ≫ H and so the mainsource of friction is dissipation into fermionic modes ratherthan Hubble expansion. The field’s motion will then beoverdamped for ϒ≳ jmϕj, where m2

ϕ ¼ V 00ðϕÞ, and under-damped otherwise. For relativistic fermions, close tothe origin we have m2

ϕ ≃ α2ðT2 − T2cÞ, while for ϕ ¼ �v

the field mass is m2ϕ ¼ α2ð2T2

c þ T2Þ. This means that thefield’s trajectory from the symmetric to the symmetry-breaking minimum will be overdamped if during thetransition the temperature is above 0.05gTc=

ffiffiffiffiffiffiffiNF

p, which

is parametrically below the critical value. This implies thatinstead of oscillating about ϕ ¼ �v, the field will smoothlyevolve towards this value. When the motion is overdamped,the scalar field equation reduces to a slow-roll equation ofthe form

3Hð1þQÞ _ϕ≃ −V 0ðϕÞ; ð26Þ

where Q≡ϒ=3H, such that in the radiation era Q ∼8.5

ffiffiffiffiffiffiffiNF

pMP=T ≫ 1 as argued above and the field’s evo-

lution occurs in a strong dissipation regime. Furthermore,the field may remain close to the origin and mimic acosmological constant if its energy density does not varysignificantly within a Hubble time:

1

H

_ρϕρϕ

≃ −3ð1þQÞ_ϕ2

V≃ −2

Ωϕϵϕ1þQ

≪ 1; ð27Þ

where ϵϕ ¼ ðM2P=2ÞðV 0ðϕÞ=VðϕÞÞ2 is the slow-roll param-

eter typically considered in slow-roll inflationary modelsand Ωϕ ¼ ρϕ=ρtotal is the Higgs field relative abundance.

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This condition essentially ensures that the Higgs field doesnot dissipate a significant fraction of its energy density intothe heat bath on cosmological time scales, thus sustaining acosmological constant–like behavior for small kineticenergy. Note also that the condition above reduces to theslow-roll condition ϵϕ ≪ 1 in nondissipative (cold) infla-tionary models, where the scalar field is the dominantcomponent, and to the slow-roll condition ϵϕ ≪ 1þQ indissipative (warm) inflation scenarios. It moreover showsthat a constant energy density is easier to maintain when thefield is subdominant, Ωϕ < 1.Close to the origin we find ϵϕ ≃ 8ðMP=vÞ2ðϕ=vÞ2,

which can be small if the field is very close to the origin.Inflation is, however, hard to maintain in the absence ofdissipation with this type of “hill-top" potential since thecurvature parameter ηϕ ¼ M2

PV00=V ≃ −4M2

P=v2 is too

large unless v ≳MP. Dissipation into the heat bath alle-viates this constraint by overdamping the field’s motionas shown above. As first shown in [60,62], it is hard toobtain a very long period of inflation with the dissipationcoefficient in Eq. (24), in particular the 50–60 e-foldsrequired to solve the horizon and flatness problems, sincethe fermion mass increases as the field moves towards theminimum and eventually dissipation becomes Boltzmannsuppressed. We note, however, that for supersymmetricmodels in the low-temperature regime, where dissipation isdominantly mediated by low-momentum scalar fieldmodes, fully successful models of warm inflation havebeen developed (see e.g. [22,31,46,47]). In such scenariosdissipation can sustain both the slow-roll dynamics ofthe inflaton field and the temperature of the radiation bathfor a sufficiently long period. Nevertheless, the dissipationcoefficient in Eq. (24) can sufficiently overdamp the field’smotion to allow for a few e-folds of inflation which,analogously to thermal inflation, can dilute dangerousrelics generated prior to the phase transition. The slow-roll equation (26) can be solved when the field is close tothe origin in a radiation-dominated universe, yielding

ϕ ∝ exp

�0.1α2

N3=2F

MP

T

��Tc

T

�2

− 3

��; ð28Þ

so that the field increases exponentially below the criticaltemperature. Since ϵϕΩϕ=Q ∝ ϕ2=T3 in this case, slow rollcan only be maintained for a finite period of time. TheHiggs field may become dominant at a temperatureTϕ ≃ g−1=4� αTc=

ffiffiffiλ

p, so if the scalar self coupling is

sufficiently small the field will overcome the radiationbefore the end of the slow-roll regime and induce a periodof inflation. In Fig. 2 we show a numerical example of aphase transition close to the GUT scale where a short periodof inflationary expansion occurs.It is clear in this figure that the field’s motion is always

overdamped, exhibiting no field oscillation, even though

slow roll is only maintained for a finite period, withinflation lasting in this example for ∼1.5 e-folds. The fieldthen evolves quickly to the symmetry-breaking minimum,which is actually time dependent until the field’s thermalmass becomes exponentially suppressed. We note that thistransition is fast in terms of the cosmological Hubble time,although still adiabatic from the microphysical perspective.We emphasize that dissipation prevents the field fromoscillating about the minimum, as opposed to what iscommonly considered in phase transitions when this effectis not taken into account. Therefore, in the presence ofdissipation, the Higgs field will not behave as pressurelessmatter after the phase transition.During the Higgs-dominated phase, the radiation density

is diluted exponentially by the accelerated expansion untilthe end of the slow-roll regime. With the increase in thefield’s velocity, the dissipative source term in the radiationevolution equation (25) grows substantially, allowingradiation to once more become the dominant component.We note that in warm inflation models where 50–60 e-foldsof inflationary expansion can be sustained, generically inthe low-T rather than the high-T regime considered here, ρRtypically reaches a quasisteady evolution with the dissipa-tive source term balancing the Hubble dilution effect. In theexample shown above, inflation does not last sufficientlylong for this quasiequilibrium to be reached, with firstdilution and then dissipation playing a dominant role in theradiation evolution.A crucial point to emphasize is that the expansion history

can be significantly modified even if inflation does notoccur, i.e. in parametric regimes where slow roll cannot besustained until the field can dominate. On the one hand,

FIG. 2 (color online). Evolution of the field and radiationenergy density during a phase transition (solid lines), for a casewhere the field comes to dominate the energy density. The dashedred line shows the evolution of the radiation energy density in theabsence of a phase transition. In this example, λ ¼ 0.01,NF ¼ 10, v ¼ 1015 GeV, g2 ¼ 1=NF and h ¼ 0.1. Time hereis shown in units of the scale 1015 GeV. The vertical dashed lineindicates the time when the critical temperature is reached.

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when the field is slowly rolling, its energy density increasesthe expansion rate, therefore diluting the ambient radiationmore quickly even if it is subdominant. On the other hand,once slow roll is over and the field quickly settles into thesymmetry-breaking minimum, it can dissipate a significantpart of its energy density into the heat bath. This isillustrated in Fig. 3, where slow roll ends just before thescalar field’s abundance becomes comparable to theheat bath.This figure clearly shows that the most significant effects

occur at the end of the slow-roll regime, when the Higgsfield’s relative abundance is maximal, leading first to adilution and then to an increase in the radiation energydensity. The latter eventually relaxes to the value it would

have in the absence of a phase transition, since the relationρRðtÞ is an attractor of the Friedmann equation in aradiation-dominated universe. One can solve the equationof motion for radiation in the absence of a dissipativesource to find

ρRðtÞ ¼3m2

p

4t2

�1þ 1

t

� ffiffiffi3

pmp

2ρ1=2R0

− t0

��−2: ð29Þ

At sufficiently large t the radiation becomes insensitive toits initial value and shows an attractor behavior tendingtowards ρRðtÞ ¼ 3m2

p=4t2. This explains why the radiationapproaches the standard evolution given by the dotted redline in Figs. 2 and 3 at late times after dissipation (particleproduction) becomes irrelevant. However, the increase inthe Hubble parameter during the phase transition makes theUniverse expand by a larger factor than in the standardradiation-domination scenario. In Fig. 4 we show theevolution of the Hubble parameter and radiation energydensity relative to a standard radiation-dominated universe,for different numbers of fermion species. As one can easilyconclude, increasing NF enhances the effect of dissipationand the relative change of H and ρR.The additional expansion will have a diluting effect

on any decoupled particle species, for which the numberdensity redshifts as n ∝ a−3. This includes e.g. topologicalrelics such as monopoles or thermal relics such as grav-itinos generated prior to the phase transition. Consideringan initial time ti before the critical temperature is reachedand a final time tf after the field has settled into thesymmetry-breaking minimum, we have for a genericdecoupled species

nf ¼ ni exp�−3

Ztf

ti

Hdt0�: ð30Þ

FIG. 3 (color online). Evolution of the field and radiationenergy density during a phase transition (solid lines), for a casewhere the field never dominates the energy density. The dashedred line shows the evolution of the radiation energy density in theabsence of a phase transition. In this example, λ ¼ 0.2, NF ¼ 1,v ¼ 1015 GeV, g2 ¼ 1=NF and h ¼ 0.1. Time here is shown inunits of the scale 1015 GeV. The vertical dashed line indicates thetime when the critical temperature is reached.

FIG. 4 (color online). In the left- (right-) handed plot we show the evolution of the radiation energy density (Hubble parameter)compared to that of a standard radiation-dominated universe. In this example, λ ¼ 0.1, v ¼ 1015 GeV, g2 ¼ 1=NF and h2 ¼ 0.1, Timehere is shown in units of the scale, 1015 GeV.

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Assuming no changes in the number of relativistic degreesof freedom, the entropy density of the heat bath before andafter the phase transition is related by sf=si ¼ ðTf=TiÞ3.This implies that the number density-to-entropy ratio of thedecoupled species is diluted by a factor,

nf=sfni=si

¼ exp ½−3 R tfti Hdt0�

ðTf=TiÞ3: ð31Þ

In Fig. 5 we show numerical results for this dilution factoras a function of the number of fermion species, showingthat stronger dissipative effects lead to a more significantdilution of dangerous relics, by enhancing either themaximum value of Ωϕ attained or the duration of the lateperiod of warm inflation. For example, observationalconstraints on the abundance of GUT monopoles requireat least nM=s≲ 10−11 [63,64], so it is unlikely that a singlephase transition subsequent to monopole formation canyield the required dilution factor unless a very large numberof dissipative channels is involved. Even if a completedilution cannot be achieved, this may, for example, alle-viate the bounds on the reheating temperature after inflationconcerning the overproduction of gravitinos. In particular,since Ω3=2 ∝ TR (see e.g. [65]), bounds on TR will increaseby the inverse of the dilution factor in Eq. (31).Furthermore, the cumulative effect of several differentstages of symmetry breaking may potentially result in asignificant dilution factor that should be taken into account.In summary, we have shown that dissipative effects

during a cosmological phase transition may have a sig-nificant effect on the cosmic history. By overdamping themotion of the associated Higgs field, dissipation not onlyprevents oscillations about the symmetry-breaking mini-mum but also leads to a period of slow roll and potentiallylate-time warm inflation. The energy density in the fieldand the entropy produced by dissipative effects will also

generically increase the amount of Hubble expansionduring the phase transition and parametrically dilute theabundance of frozen relics.One or more short periods of late-time warm inflation

during phase transitions could have significant observa-tional effects. On the one hand, their existence implies thatthe main period of inflation can be considerably shorterthan the overall 50–60 e-folds of accelerated expansionrequired by the observed flatness and homogeneity of theUniverse. This will therefore change observational predic-tions for large scales, along the lines suggested in [66] forthe case of thermal inflation. On the other hand, small scaleperturbations will be generated during these periods,although they should be well within the horizon todayand hence potentially too damped to be studied in galaxysurveys or CMB observations. Although this requiresfurther inspection and a detailed study that is outside thescope of this work, we nevertheless emphasize that dis-sipation will modify the evolution of fluctuations, thereforeyielding distinct observational predictions from a periodof thermal inflation. Since both thermal and dissipative(warm) inflation may occur within the same phase tran-sition, it would be interesting to explore the combinedeffects of these two types of inflationary expansion on thespectrum of cosmological perturbations.

IV. DISSIPATIVE BARYOGENESISAND LEPTOGENESIS

As we have seen in the previous section, dissipation mayhave a significant effect in the dynamics of a cosmologicalscalar field in the process of spontaneous symmetry break-ing. Significant effects arose in this case when the fieldbecame a non-negligible component of the energy balancein the Universe, either itself increasing the Hubble rate orleading to a significant entropy production. In this section,we consider an effect of dissipation that may occur evenwhen the dissipating scalar field carries a very smallfraction of the energy in the Universe and plays asubdominant role in entropy production.Dissipation leads to the production of particles within

the heat bath to which a dynamical scalar is coupled to,continuously disturbing its equilibrium. The degrees offreedom within the heat bath will a priori include the SMparticles and their antiparticles, as well as potentially darkmatter particles and other beyond the SM species. The rateat which each particle species is produced is related to itsfractional contribution to the dissipation coefficient, asexplicitly shown in [67]. It is then natural to envisagescenarios where particles and antiparticles are producedat different rates by a dissipating scalar field, necessarilyinvolving interactions that violate baryon/lepton number aswell as the C and charge parity (CP) symmetries, accordingto the conditions first established by Sakharov [68]. Thiswas first explored in the context of warm inflation in amechanism dubbed warm baryogenesis [35], where the

FIG. 5 (color online). Dilution factor for frozen relics during aphase transition as a function of the number of fermionscoupled to the Higgs field, for λ ¼ 10−2, v ¼ 1015 GeV, g2 ¼1=NF and h ¼ 0.1.

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same interactions responsible for damping the inflaton’smotion and sustaining a radiation bath during inflationwere shown to yield a significant baryon asymmetry,parametrically within the observed window.Here, we show that dissipative baryogenesis is a much

more general mechanism that may occur in the dynamics ofany cosmological scalar field with nonequilibrium dissipa-tive dynamics and interactions satisfying the Sakharovconditions. We illustrate this by looking at a concreteexample based on the interactions employed in standardthermal leptogenesis scenarios with right-handed neutrinosthat is naturally motivated within GUT models. Althoughdissipative baryo/leptogenesis will occur in several differentdynamical regimes, we focus on low-temperature dissipativemodels to explicitly show that the production of a leptonasymmetry does not require temperatures above the right-handed neutrino mass threshold as in the standard thermalscenarios. Our example further shows that no symmetriesneed to be restored in the early Universe to generate theobserved baryon asymmetry, thus avoiding the severalpotential cosmological problems that this may cause.We first consider the relevant particle physics inter-

actions and describe how they lead to dissipative effectsthat may produce more particles than their antiparticles, andafterwards describe the dynamics of dissipative baryo-genesis in the radiation-dominated era.

A. Interactions and dissipative particle production rates

Leptogenesis is amongst the most popular models for thegeneration of a cosmic baryon asymmetry [69–74]. In thesimplest models, it is based on the out-of-equilibriumdecays of heavy right-handed Majorana neutrinos, whichviolate lepton number as well as C and CP. The resultinglepton asymmetry is later on converted into a baryonasymmetry by electroweak sphaleron processes [75], whichconserve B − L but not the two global charges independ-ently. Heavy right-handed neutrino singlets are the simplestaddition to the SM particle content, yielding light neutrinomasses through the seesaw mechanism, thereby providingan interesting connection between cosmology and low-energy particle physics.Right-handed neutrinos also fit nicely within the 16

fundamental representation of the SOð10Þ GUT gaugegroup and their large Majorana mass required by theseesaw mechanism can in this case be generated by thevev of a Higgs field in the 126 representation [76]. It is thusnatural to consider the cosmological dynamics of this scalarfield, which to our knowledge remains unexplored, includ-ing in particular the dissipative effects associated with itscouplings to right-handed neutrinos. We then consider asupersymmetric model where the relevant interactions areencoded in the superpotential,

W ¼ 1

2gaΦNc

aNca þ yaiNc

aHuLi þ fðΦÞ; ð32Þ

which involves the right-handed neutrino superfields, Na,as well as the SM lepton and Higgs doublet superfields, LiandHu, respectively. We consider three neutrino and leptongenerations denoted by the indices a and i and note thatSUð2Þ gauge indices are implicit in the superpotential. Thechiral superfield Φ can be identified with the scalardirection within the 126 representation of SOð10Þ thatgives a Majorana mass to the right-handed neutrinos, asdiscussed above, or more generally as a SM singlet withself interactions encoded in the analytic function fðΦÞ.Without loss of generality, we take its scalar vev as a realfield hΦi ¼ ϕ=

ffiffiffi2

p.

Dissipation of the scalar field’s energy will in this caseproceed through the excitation of the right-handed neu-trinos and their scalar superpartners in the cosmic heat bathand their subsequent decay into the MSSM (s)leptons andHiggs(inos). The cosmological evolution of the ϕ field willdepend on its potential, given by jf0ðϕÞj2, and crucially onits behavior during inflation. As anticipated above, we willbe mainly interested in studying the regime where right-handed neutrinos are too heavy to be thermally producedand therefore standard leptogenesis scenarios are ineffi-cient. This is natural in scenarios where the field is light andhence overdamped during inflation, either remaining frozenat some potentially large initial value or driven towards alarge vev by de Sitter fluctuations. In the low-temperatureregime where the reheat temperature after inflation is belowthe right-handed neutrino mass threshold, dissipationproceeds through the excitation of virtual modes in theheat bath as discussed earlier in this work. Scalar modes, inthis case the right-handed sneutrinos decaying dominantlyinto sleptons and Higgs bosons, yield the leading contri-bution to the dissipation coefficient [47], which is givenapproximately by

ϒ ¼ CϕT3

ϕ2; Cϕ ≃ 1

16π

Xa;i

yaiy�ai: ð33Þ

This coefficient therefore determines the overall entropyproduction rate in the form of MSSM particles producedin the thermal bath by the decays of the virtual right-handed sneutrinos. These decays violate lepton number,since the Majorana mass term precludes a consistentassignment of L to Nc

a, and may also violate C and CP ifthe Yukawa coupling matrix has nontrivial phases, whichis possible for at least three matter generations. If this isthe case then out-of-equilibrium dissipation will naturallyinduce an overabundance of sleptons over antisleptons inthe heat bath (or vice versa, although we assume this to bethe case).The rate at which sleptons and antileptons are produced

can be computed from the imaginary part of their selfenergies, following the generic procedure first describedin [67]:

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_nðpÞ ¼ Im

�2

Zt

−∞dt2

e−iωpðt−t2Þ

2ωpΣ21ðp; t; t2Þ

�: ð34Þ

Integrating over the 3-momentum p and summing over theenergies of all the light particle species yields a source termfor the radiation energy density ϒ _ϕ2 corresponding to thedissipation coefficient in Eq. (33). We are, however,interested in the difference between the slepton and anti-slepton production rates, which as we will show is asubleading effect compared to the overall dissipativeentropy production. To compute the slepton self energieswe first consider the relevant scalar and Yukawa inter-actions resulting from the superpotential in Eq. (32), whichare given by

Ls ¼g2a2ϕ2j ~Naj2 þ

gaϕffiffiffi2

p yaj ~N�ahu~lj þ H:c: ð35Þ

LY ¼ 1

2yaið ~Nc

a~huPLli þ ~Nc

aliPL~hþ huliPLNa þ huNaPLli

þ ~li ~huPLNa þ ~liNaPL~hu þH:c:Þ: ð36Þ

In Fig. 6 we show the leading one- and two-loop diagramscontributing to the slepton self energies. We note that, eventhough dissipation is dominated by scalar modes, at two-loop order fermions and scalars will give comparablecontributions to the slepton self energy. It is also interestingto note that even though the final asymmetry is independentof the lepton number assignment chosen for the right-handed neutrinos, this choice determines which diagramsactually exhibit L-violation. For example, the Yukawasector always violates L, whereas scalar interactionspreserve it for LðNc

aÞ ¼ 1. In scenarios where B − L is aspontaneously broken gauge symmetry, the Nc

a superfield

will have lepton number −1 and L is violated by both typesof interactions.Analogously to standard leptogenesis scenarios, CP

violation arises only through the interference betweenthe leading and next-to-leading diagrams. The leadingdiagram corresponds in this case to the top-left diagramin Fig. 6, the imaginary part of which yields the (tree-level)decay of the right-handed sneutrinos. The slepton selfenergies can be computed using standard thermal fieldtheory techniques and we refer the reader to [35,67] formore technical details. Slepton and antislepton self energiesare related by charge conjugation and we obtain for thedifference between the self energies, to leading order,

ΔΣ~li21 ¼

3

16π

Xa;b

Xj

Zdp

ZdkCðp; kÞ

×

�Tmb

�4 mb

maImðybiy�ajybjy�aiÞ; ð37Þ

where ma ¼ gaϕ=ffiffiffi2

pare the right-handed sneutrino

masses to leading order, assuming the MSSM Higgs andsleptons have zero or at least negligible expectation valuesat this stage in the cosmological evolution. We note thatonce the sum over all heavy sneutrino and light fieldgenerations is performed, only the diagrams involvingfermionic propagators contribute to the self-energy differ-ence above. The factor Cðp; kÞ is common to the selfenergy of all the different particle species and thereforedetermines the overall dissipation coefficient given above.Adding the self energies of all particle species we obtain toleading order the following overall radiation productionrate:

ΣR21 ¼ 8

Xa

Xj

Zdp

ZdkCðp; kÞyajy�aj: ð38Þ

FIG. 6. The one-loop and two-loop diagrams contributing to the slepton self energy. Black circles indicate background field (i.e. right-handed sneutrino mass) insertions.

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The relative rate at which a lepton asymmetry is producedby dissipation can then be obtained by taking the quotientof Eqs. (37) and (38), yielding

rL ≡ _nL_ndR

≃ 3

64π

1Paðyy†Þaa

Xa≠b

�Tmb

�4 mb

maIm½ðyy†Þ2ba�;

ð39Þwhere the sum over light fields running in the loop isimplicit. Note that, as in thermal leptogenesis, a non-vanishing asymmetry can only be produced if at least twoof the right-handed sneutrinos are nondegenerate, thusrequiring distinct ga couplings in the superpotential (32).A couple of important properties of the asymmetry

production rate should be emphasized. First, the dissipationcoefficient in Eq. (33) is independent of the ga couplings toleading order, so that all three right-handed sneutrinospecies will be virtually excited by the motion of the ϕfield and contribute to the lepton asymmetry. This is incontrast to thermal scenarios, where the out-of-equilibriumdecay of the lightest right-handed (s)neutrino will give adominant contribution. Second, the asymmetry productionrate is suppressed by ðT=maÞ4 ≪ 1, which is associatedwith the fact that the right-handed sneutrinos are onlyvirtually excited, as opposed to thermal leptogenesis sce-narios. This means that while in the latter mechanism onemust consider small couplings and CP violating phases toyield the observed baryon asymmetry; in dissipative lepto-genesis a small baryon-to-entropy ratio can result solelyfrom the low-temperature suppression. We note that theleading scalar loop diagrams contributing to the asymmetryare only suppressed by a factor ðT=maÞ2, but as mentionedabove their overall contribution cancels out when summingover the different generations. This is a specific feature ofthe interaction structure considered in leptogenesis, with asingle type of decay channel for the heavy right-handedsneutrinos, so that in more general models of dissipativebaryogenesis, such as the one considered in [35], theasymmetry production rate will be larger.The light neutrino mass hierarchy inferred from exper-

imental bounds motivates considering a hierarchicalstructure in the right-handed neutrino sector as well, e.g.g1 ≪ g2 ≪ g3. In this case the asymmetry production ratereduces to

rL ≃ 3

64π

�Tm1

�4 1P

aðyy†ÞaaXNR

a≠1

�m1

ma

�Im½ðyy†Þ21a�;

ð40Þwhich is suppressed relative to its thermal leptogenesiscounterpart by a factor ð1=8ÞðT=m1Þ4, as well as the factthat in the latter case only the lightest right-handed neutrinocontributes to the factor

Paðyy†Þaa in the denominator

which corresponds to the overall entropy production rate.

To simplify our dynamical analysis of dissipative lepto-genesis, we collect all couplings and mass differences intoan effective parameter ϵ, such that

rL ≃ ϵy2�Tm1

�4

; ð41Þ

where we assumed that the Yukawa couplings have roughlythe same magnitude y.

B. Dynamics of the lepton asymmetry generation

Having determined the rate at which lepton number isproduced by dissipation, we now consider the dynamics ofthe scalar field ϕ, which is coupled to the evolution of theoverall entropy and lepton number density via the system ofequations

ϕþ ð3H þϒÞ _ϕþ V 0ðϕÞ ¼ 0; ð42Þ

_sþ 3Hs ¼ ϒ _ϕ2

T; _nL þ 3HnL ¼ 45ζð3Þ

2π4gLg�

ϒ _ϕ2

TrL;

ð43Þ

where gL is the number of relativistic degrees of freedomwith nonvanishing lepton number, for which we will takethe MSSM value gL ¼ 33.75 as a reference, as well asthe associated g� ¼ 228.75 for the overall number ofrelativistic species. We consider the evolution in thelow-temperature dissipative regime, where the dissipationcoefficient takes the form in Eq. (33). We assume that theYukawa couplings have roughly the same magnitude forall three generations, such that Cϕ ≃ 9y2=16π, althoughthis assumption is not crucial for our subsequent analysis.The lepton number density is sourced, as computed in theprevious section, by a fraction rLϒ of the overall dis-sipation coefficient.We are interested in the evolution of the field ϕ in a

regime where it has a large vev, such that right-handed (s)neutrinos have a large mass and cannot be produced onshell. This implies that, as opposed to the exampleconsidered in the previous section, we assume that thereis no symmetry restoration after inflation. As a concreteexample, we take the simple symmetry-breaking potentialof the previous section, given in Eq. (22), although thermalmass corrections will always be Boltzmann suppressed inthe regime that we are interested in exploring.We note that if the field ϕ comes to dominate over the

radiation energy density, or at least attains a significantrelative abundance, then a sizeable lepton asymmetry canbe produced, which is the case of the warm baryogenesismechanism during inflation [35]. We show, however, thatthe observed baryon asymmetry can also be produced whenthe field is subdominant and dissipation does not contributesignificantly to the overall entropy of the Universe.

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It is convenient to express the lepton number densityin terms of the lepton-to-entropy ratio, YL ≡ nL=s, whichbecomes constant once a lepton asymmetry stops beingefficiently produced. We then have

_YL ¼ CLT3

ϕ6_ϕ2; CL ¼ 90ζð3Þ

π4gLg�

Cϕ

Csϵy2

g4; ð44Þ

where Cs ¼ 2π2g�=45. The evolution of YL will then bedetermined by the dynamics of the ϕ field. We assume thefield’s self coupling λ is sufficiently small for it to beoverdamped during inflation, m2

ϕ ¼ λ2ð3ϕ2 − v2Þ ≪ H2inf ,

where Hinf ≲ 1014 GeV from the recent CMB upperbounds on the tensor-to-scalar ratio obtained by thePlanck satellite [77]. De Sitter fluctuations will then leadto a distribution of field values ϕi ≲MP in different patchesof the inflationary universe at the start of the radiation era,which will typically be displaced from the minimum of thepotential at jϕj ¼ v.As it evolves towards the minimum, the field will feel the

effects of both Hubble damping and dissipative friction.The latter will play a significant role for

Q ¼ ϒ3H

≃ffiffiffiffiffi10

π2

rCϕffiffiffiffiffig�

p TMP

ϕ2≳ 1; ð45Þ

where we have used the standard relation between theHubble rate and the ambient temperature in a radiation-dominated universe. On the one hand, if the field rollstowards the minimum from jϕij < v, Q will necessarilydecrease in time and so dissipation can at most have asignificant effect during the earlier stages of the evolution.On the other hand, for jϕij > v, Qmay increase as the fieldvalue decreases, in particular if it overshoots the minimumand attains a small value during the first oscillation. In anycase dissipation can only have a transient effect, sinceasymptotically the field will settle at the minimum and Qwill decrease with the temperature.For simplicity, we focus on scenarios where dissipation

plays no significant role in the field dynamics. This is,in particular, the case for a large field vev v ≲MP andinitial displacements Δϕi ¼ jϕi − vj≲ v, for whichQ≲ T=MP ≪ 1. In the standard seesaw mechanism theright-handed (s)neutrino masses are related to the lightneutrino masses via

m ~N ¼ gvffiffiffi2

p ≃ 1015y2�0.1 eVmν

�GeV; ð46Þ

so that v ≲MP implies g≳ 10−3y2ð0.1 eV=mνÞ. Underthese conditions the field oscillations are well described by

ϕ≃ vþ Δϕi

�tti

�−3=4

cos ðmϕðt − tiÞÞ; ð47Þ

where mϕ ¼ ffiffiffi2

pλv is the field mass at the minimum,

Δϕi ≡ ϕi − v and ti ¼ ð2HiÞ−1 ∼ ð2mϕÞ−1 is the time atwhich the field becomes underdamped and effectivelystarts oscillating. We may then substitute this intoEq. (44) to estimate the lepton-to-entropy ratio producedby dissipation as the field oscillates. We note that sincethe field velocity is small before the onset of oscillations,no significant lepton number will be produced until thefield becomes underdamped. Taking the average fieldvalue hϕi ¼ v and the average field velocity h _ϕ2i≃m2

ϕΔϕ2i ðt=tiÞ−3=2=2, we can then integrate Eq. (44) from

t ¼ ti to obtain asymptotically

YL ≃ CL

4g3=4�

�45

2π2

�3=4

ðMPtiÞ3=2m2

ϕΔϕ2i

v61

t2i: ð48Þ

For g� ¼ 228.75 and gL ¼ 33.75 in the MSSM, and takinginto account the relation between the asymptotic baryonand lepton numbers after conversion by sphaleron proc-esses, Bf ¼ −ð8=23ÞLi [74], this yields

ηs ≃ 0.1

�mν

0.1 eV

�2�

m ~N

1015 GeV

�1=2

�Δϕi

v

�2�mϕ

m ~N

�5=2

ϵ:

ð49Þ

The baryon asymmetry thus depends parametrically on theratio of the field and sneutrino masses, mϕ=m ~N ¼ 2λ=g.Adiabaticity of the dissipative process requires the field tomove slowly compared to the sneutrino decay rate, _ϕ=ϕ∼mϕ ≪ Γ ~N ≃ y2m ~N=16π. Additionally, for the asymmetryto be produced below the sneutrino mass threshold, werequire T < m ~N at the onset of field oscillations, H ∼mϕ.This then implies mϕ=m ~N ≲m ~N=MP, which is typically astronger constraint than adiabaticity of the dissipativeprocesses. Saturating this bound, we obtain for the finalbaryon asymmetry

ηs ≃ 10−10�

mν

0.1 eV

�2�

m ~N

2 × 1015 GeV

�3�Δϕi

v

�2�

ϵ

0.05

�:

ð50Þ

We thus see that the observed baryon asymmetry, ηs ∼10−10 [78], can be obtained through adiabatic dissipationfor sneutrino masses close to the GUT scale, whichgenerate neutrino masses in the range suggested byatmospheric neutrino oscillations for Oð1Þ Yukawa cou-plings [79]. As opposed to standard leptogenesis models,the amount of CP violation, parametrized by ϵ, need not bevery small in this case since the produced baryon asym-metry is naturally small. We note that Eq. (50) is anestimate that is accurate up to Oð1Þ factors, since the onsetof field oscillations does not occur exactly for H ¼ mϕ, so

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the above values for the masses should be taken only asreference values.The exact value of the produced asymmetry can be

computed numerically, and in Figs. 7 and 8 we giveexamples for the numerical evolution of the field ϕ andthe asymmetry nB=s in the regime considered above, fordifferent values of the field mass parametrized by the selfcoupling λ. In all examples shown, dissipation has asubdominant effect on the field evolution, as discussedabove, and its main effect is the production of a baryonasymmetry. We have checked in all cases that the adiabaticcondition is satisfied and that the temperature is below thesneutrino mass threshold at the onset of field oscillations.We thus conclude that the produced asymmetry can have

a range of values both below and above the observationalwindow. Most of the lepton number is produced in the firstfew oscillations of the field, where the field velocity, andhence the dissipative lepton source, is larger, with the

lepton-to-entropy ratio stabilizing within a few oscillationperiods.We note that, even though the adiabatic dissipation

coefficient decreases with the temperature, and hencebecomes negligible at late times, the full dissipationcoefficient includes a zero-temperature part that corre-sponds to the standard decay width for an oscillating field[67]. This corresponds in the present scenario to the four-body decay of the ϕ field into Higgs and slepton pairsmediated by virtual right-handed sneutrinos, since thelatter’s on-shell production is kinematically forbidden.As shown in [67], this contribution is suppressed byðmϕ=TÞ3, as well as numerical factors, with respect tothe adiabatic component. We may thus safely neglect thiscontribution in computing the lepton asymmetry, which isproduced whenmϕ ∼H ≪ T, bearing nevertheless in mindthat this will lead to the decay of the ϕ field after it becomesnonrelativistic at late times.

FIG. 8 (color online). Numerical results for the evolution of the field ϕ and the baryon asymmetry ηs ¼ nB=s, starting from a fieldvalue above the minimum at v ¼ MP. The quartic self coupling λ ¼ 1 × 10−7; 2 × 10−7; 4 × 10−7 for the blue, black and green curves,respectively. We have taken g ¼ 10−3, y ¼ 3 and ϵ ¼ 1=64π in all cases. Time is given in Planck units.

FIG. 7 (color online). Numerical results for the evolution of the field ϕ and the baryon asymmetry ηs ¼ nB=s, starting from a fieldvalue below the minimum at v ¼ 1018 GeV. The quartic self coupling λ ¼ 1 × 10−8; 2 × 10−8; 4 × 10−8 for the blue, black and greencurves, respectively. We have taken g ¼ 10−3, y ¼ 3 and ϵ ¼ 1=64π in all cases. Time is given in Planck units.

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In the particular model of leptogenesis that we haveconsidered, a lepton (and hence baryon) asymmetry isproduced by the dynamical evolution of a scalar SM singletthat determines the Majorana mass of right-handed (s)neutrinos. This is, however, a much more general result anddissipative baryogenesis should occur in any scenariowhere fields whose decay violates the B/L and C, CPsymmetries are coupled to (and acquire mass from) adynamical scalar field, including e.g. the SUð5Þ modelconsidered in Sec. II. Depending on the field masses andcouplings, the observed baryon asymmetry may be entirelyproduced by off-shell dissipative effects, with no need fortemperatures above the B-violating field mass threshold.This may then avoid symmetry restoration in the earlyUniverse and the production of dangerous thermal relicsduring the associated phase transitions. In addition, dis-sipative baryogenesis generically yields potentially observ-able signatures, as we describe below.

C. Isocurvature perturbations

As obtained above, the baryon asymmetry that resultsfrom dissipative processes will depend on the initial fielddisplacement from the true minimum of its potential. If, aswe assumed earlier, the field is light during inflation, wethen expect superhorizon quantum fluctuations hδϕ2

i i ¼ðHinf=2πÞ2 in the initial field value. These will then result influctuations in the final baryon-to-entropy ratio and hencebaryon isocurvature modes that can be tested with CMBobservations.This is also a feature of the warm baryogenesis scenario

during (warm) inflation [35], where both inflaton andtemperature fluctuations generate baryon isocurvaturemodes. The main difference to the case analyzed in thiswork resides, first, in the fact that the field responsible forproducing the baryon asymmetry never dominates theenergy balance in the Universe. Consequently, the resultingbaryon isocurvature modes will be uncorrelated with themain (adiabatic) curvature perturbations sourced by theinflaton. Second, the baryon asymmetry is produced atthe onset of field oscillations rather than in a slow-rollregime. Since this occurs when H ∼mϕ, fluctuations inthe ambient temperature will only delay or expedite theproduction of the baryon asymmetry, but they do notchange its final value. From Eq. (50), we have that

SB ¼ δηsηs

¼ 2δϕi

Δϕi; ð51Þ

such that we can write the relative contribution ofuncorrelated baryon isocurvature modes to the CMBspectrum as

B2B ¼ S2B

P2ζ

≃ r2

�Mp

Δϕi

�2

; ð52Þ

where P2ζ ≃ 2 × 10−9 is the amplitude of the adiabatic

curvature perturbation spectrum [77] and r is the tensor-to-scalar ratio. This gives a contribution to the total matterisocurvature power spectrum that is suppressed by therelative abundance of baryons ðΩb=ΩmÞ2B2

B ≃ 0.03B2B.

From the constraints posed by the Planck satellite onuncorrelated cold dark matter (CDM) isocurvature modeswith a scale invariant spectrum at a comoving wavenumber klow ¼ 0.002 Mpc−1 [77], we deduce the boundBB ≲ 1.03. The result above satisfies this bound for initialfield displacements:

Δϕi > 0.68ffiffiffir

pMP: ð53Þ

If the tensor-to-scalar ratio is close to the current upperbound r≲ 0.1 [77], this requires field displacementsΔϕi ≳ 0.2MP. Although the value of the baryon asym-metry does not depend directly on the actual value ofthe field displacement, but rather on the ratio Δϕi=v, wehave seen above that the observed baryon asymmetry canbe entirely produced by off-shell dissipative effects forΔϕi ∼ v ∼MP. The bound above is thus consistent withthe generation of the observed baryon asymmetry. On theother hand, low-scale inflationary models with r ≪ 0.1are consistent with initial field displacements parametri-cally below the Planck scale.We note, however, that in supergravity models scalar

fields may acquire masses parametrically close to theHubble scale during inflation, and hence be driven to alocal minimum that does not necessarily coincide with thelow-energy global minimum [80]. In this case dissipationmay also produce a baryon/lepton asymmetry as the fieldrolls towards the true minimum after inflation, althoughfield masses of the order of the Hubble scale may somewhatchange the dynamics. In these scenarios there will be,however, no significant field fluctuations on superhorizonscales, which makes them less appealing from the obser-vational point of view.Evidence for baryon isocurvature modes will never-

theless constitute a strong hint for dissipative baryogenesis,which is thus a testable mechanism. In addition, theparticular case of dissipative leptogenesis considered abovecan be related to low-energy neutrino phenomenology, thusyielding two independent potential ways of probing theproduction of a baryon asymmetry.

V. CONCLUSION

Scalar fields are ubiquitous in the best-motivated exten-sions of the standard model of particle physics and theirdynamics has in most cases a very significant cosmologicalimpact. Since they generically interact with other matterand gauge degrees of freedom, dissipative effects are acrucial feature determining how scalar fields evolve in thecosmological heat bath. This leads to additional friction,

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entropy production and scalar field fluctuations. Theseeffects are, in the leading adiabatic approximation, fullyencoded in a single dissipation coefficient, which can becomputed from the fundamental Lagrangian defining theproperties and interactions of a given scalar field.The study of dissipative effects has so far been mostly

restricted to the early period of inflation, where dissipationmay, in fact, completely change the inflaton dynamics andthe associated generation of primordial curvature pertur-bations [22–33]. More recently, a few studies have alsobegun to explore the importance of dissipation in thedynamics of reheating after inflation [46,81–84], assumingit proceeds from a supercooled stage where dissipativedynamics plays a negligible role, and also in the dynamicsof a curvaton field [85].It was the purpose of this work to set the stage for a

much broader exploration of dissipative dynamics in theevolution of cosmological scalar fields, which in manycases only begins in the radiation-dominated era once theHubble expansion rate and the ambient temperature havedecreased sufficiently. A scalar field will typically finditself displaced from the absolute minimum of its effec-tive potential after inflation and, in evolving towards it,the field will necessarily dissipate part of its energy intothe ambient heat bath. Dissipation is thus, in particular, aninherent part of the process of spontaneous symmetrybreaking and can modify the dynamics of the severalphase transitions that may have occurred in the earlycosmic history. The evaporation of a Bose condensatemay also lead to an effective friction term along the linessuggested in [86].A natural starting point for our study was to compute, for

the first time, the dissipation coefficient inherent to differ-ent scalar fields in particle physics models of the earlyUniverse, which we hope will be useful for future studies.We have, in particular, considered the electroweak Higgsfield(s) in the SM and its minimal supersymmetric exten-sion, the scalar singlet yielding the μ term in the NMSSM,the adjoint Higgs direction breaking the GUT SUð5Þ groupto the SM, and a SM singlet giving a Majorana mass toright-handed neutrinos [embedded e.g. in SOð10Þ]. For agiven dynamical scalar, the dissipation coefficient takesdifferent forms depending on the properties of the fieldsit is coupled to, namely their mass, spin, multiplicity andcoupling constants. Dissipation proceeds genericallythrough the excitation and decay of these fields and theassociated coefficient takes different forms depending onwhether on-shell or off-shell excitation is dominant. Note,for example, that while on-shell modes resonantly enhancedissipation, their occupation numbers become Boltzmannsuppressed at temperatures below their mass thresholdand virtual modes then yield the dominant contribution.Generically, dissipation coefficients depend both on thefield value and the ambient temperature, thus constitutingdynamical quantities.

We then proceeded to explore the dynamical impact ofdissipative effects by solving the Langevin-like equationthat determines the evolution of a cosmological scalar field,Eq. (1), in different scenarios. In this work we have focusedmainly on the effects of dissipation and entropy production,although we have also briefly discussed the importanceof the dissipative noise term, which we plan to explore inmore detail in a future work. This term will be crucial, forexample, in the initial stages of a cosmic phase transition,randomly kicking the associated Higgs field away from theunstable symmetric point. Fluctuation-dissipation will thenplay an important role in the formation of topologicaldefects and also potentially in sourcing cosmic magneticfields. There are various works in the literature which makestatistical arguments for the distribution of seed magneticdomains. In a related context, statistical arguments aremade on the initial distribution of a scalar field, such asin examining the initial condition problem of inflation[87–89]. The Langevin-like Eq. (1) provides a dynamicalequation from which such distributions can be calculatedrather than just argued statistically.The friction effects associated with dissipation will slow

down a field’s evolution towards the minimum of itspotential and damp the amplitude of its oscillations aboutthe minimum. This will prolong a cosmic phase transition,potentially even completely overdamping the associatedHiggs field. One interesting outcome of our analysis is thepossibility of dissipation keeping the field in a slow-rollregime close to the symmetric value, such that it drives alate period of warm inflation. This may last for a few e-folds, which may be sufficient to dilute dangerous thermalrelics such as gravitinos or GUT monopoles formed atearlier stages. In fact, a period of thermal inflation abovethe critical temperature and a period of (dissipative) warminflation below the critical temperature can occur withinthe same phase transition. Although they have a similardilution effect, these two periods have inherently differentfield dynamics, as well as fluctuations, and we hope in thefuture to investigate more closely their potentially distinctobservational impact in the CMB and/or matter powerspectrum.We have also observed that a slow-roll period typically

ends with a Higgs field falling fast towards the symmetry-breaking minimum, in the process producing a significantamount of entropy. We have thus concluded that, even if thefield does not become dominant during the slow-roll phase,a parametric dilution of unwanted relics will still occur as acombined result of this entropy production and enhance-ment of the expansion rate during the phase transition.In this work, we have also shown that, even if

dissipation does not significantly enhance friction orentropy production, it may nevertheless lead to a smallbut nevertheless crucial effect—the generation of acosmological baryon asymmetry. This occurs due tothe out-of-equilibrium nature of dissipative processes

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when B-/L- and C-/CP-violating interactions areinvolved. We have explicitly shown that the observedbaryon asymmetry can be generated in a scenario ofdissipative leptogenesis, where a SM scalar singlet givinga Majorana mass to right-handed (s)neutrinos excites thelatter while rolling towards the minimum of its potential.This can occur at temperatures below the right-handed (s)neutrino mass threshold, with dissipation dominantlyexciting virtual modes. We expect this to be a genericfeature, also found earlier in the context of warm baryo-genesis during inflation [35], such that it is possible toproduce the observed baryon asymmetry while avoidingsymmetry restoration and subsequent production oftopological defects. Furthermore, the baryon-to-entropyratio generated through dissipation is generically fieldand temperature dependent, thus leading to baryon iso-curvature perturbations that may be probed in the nearfuture with CMB observations. If the observed asymme-try is produced by dissipation, we may thus hope to beable to test the violation of fundamental symmetries athigh energies.Our results show that going beyond the leading approxi-

mation of noninteracting adiabatic fluids can have a verysignificant impact in cosmology. Fluctuation-dissipationdynamics is present in any dynamical interacting systemand, in particular, we have explicitly shown that this isthe case for the SM Higgs field and for many otherdynamical scalars present in its extensions. Significantdissipative effects are already a feature of near-equilibriumand near-adiabatic dynamics and thus should motivatefurther exploration of this topic, not only within this regimebut also for more general nonequilibrium cosmologicalsystems.Up to now cosmology has been very successful in

explaining almost all observations through the simplemodel of a thermalized universe that is expanding. Phasetransitions have been added to this picture to reconcile itwith unified models of particle physics. There is a need topress beyond this simple picture and look with greaterdetail at the dynamics in the early Universe. The earlyUniverse is a many-body system with limited initialcondition information. A theoretical treatment of it requiresa statistical dynamical approach, which can extend on thethermal equilibrium hot big bang model. Phase transitionsand other regimes of scalar field evolution in the earlyUniverse have, up to now, only been treated classically.This paper has shown that the extension of this classicaltreatment leads to fluctuation-dissipation dynamics. In anearlier paper it was shown how fluctuation-dissipationeffects would also extend the treatment of the Universeevolution in the general hot big bang regime [42].Combined, these papers provide an extended dynamicalframework to examine key unsolved problems of the earlyUniverse.

The underlying principle behind both these papers is thesame as in the original warm inflation work—that many ofthe most fundamental quantities measured in cosmology,those associated with some underlying field dynamics, onlyprovide a coarse-grained information about that dynamicsand not necessarily about its microphysical properties. Thisis the concept that separates the warm and cold paradigmsof inflation. In cold inflation, the observables such as theindex and bispectrum are interpreted to probe preciseinformation about an underlying classical dynamics withquantum fluctuations superposed upon it. The uncertaintyin this semiclassical approach is only that associated withthe quantum mechanical uncertainty. In contrast warminflation goes further to a full statistical state where theseobservables are interpreted to provide only coarse-grainedinformation about the underlying fundamental dynamics.There can be many statistical dynamical realizations of thecoarse-grained dynamics. One common feature is thepresence of energy fluxes amongst the coarse-grained cells,often related through the underlying dynamics. The fluc-tuation-dissipation relations examined in this paper are one,perhaps most common, example of such relations, andthese can provide observable, testable, consequences.Several problems in cosmology today need to be

approached beyond the semiclassical approximation ofthermal equilibrium dynamics, such as the generation ofcurvature and isocurvature perturbations, baryogenesis,leptogenesis, generation of dark matter, origin of cosmicmagnetic fields, initial conditions of phase transitions, anddynamics during a phase transition or scalar field evolution.Adhering to just a thermal, semiclassical dynamical view-point of the underlying dynamics in the early Universe canrestrict the scope of theoretical investigation that is pos-sible, and can lead to misleading directions of interpreta-tion, such as doing elaborate model building where somesimple statistical interpretation could actually bring thepredictions in line with observation. This paper has high-lighted these points and provided a methodology that canbe used for such exploration along with several exampleapplications.

ACKNOWLEDGMENTS

We thank Rudnei Ramos for useful discussions aboutthis work. S. B. and A. B. are supported by the Science andTechnology Facilities Council (United Kingdom). J. G. R.is supported by the FCT Grant No. SFRH/BPD/85969/2012 (Portugal), as well as partially supported by GrantNo. PTDC/FIS/116625/2010 (Portugal) and the MarieCurie Grant No. NRHEP-295189-FP7-PEOPLE-2011-IRSES. J. G. R. and S. B acknowledge the hospitality ofthe Higgs Centre for Theoretical Physics of the Universityof Edinburgh and the Departamento de Física daUniversidade de Aveiro during the completion of this work.

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