the quantum phases of matter and gauge-gravity dualityqpt.physics.harvard.edu/talks/ufl13.pdfthe...

The quantum phases of matter and

gauge-gravity duality

HARVARD

University of Florida, Gainesville, February 21, 2013

Subir Sachdev

Thursday, February 21, 13

Sommerfeld-Bloch theory of metals, insulators, and superconductors:many-electron quantum states are adiabatically

connected to independent electron states


Modern phases of quantum matterNot adiabatically connected

to independent electron states:many-particle

quantum entanglement


=1⌃2

(|⇥⇤⌅ � |⇤⇥⌅)

Hydrogen atom:

Hydrogen molecule:

= _

Superposition of two electron states leads to non-local correlations between spins

Quantum Entanglement: quantum superposition with more than one particle


_



_


Einstein-Podolsky-Rosen “paradox”: Non-local correlations between observations arbitrarily far apart


1. Z2 Spin liquid in the kagome antiferromagnet

2. Superfluid-insulator transition of ultracold atoms in optical lattices:

Quantum criticality and conformal field theories

3. Holography and the quasi-normal modes of black-hole horizons

4. Strange metals: What lies beyond the horizon ?

Outline


Crystals used in collaborations: 1) Polarized Raman scattering - Lemmens, Braunschweig 2) NMR – Takashi Imai, McMaster University 3) Thermal conductivity - Behnia, Ecole Superieure 4) µSR - Keren, Technion

ZnCu(OH)6Cl2

herbertsmithite single crystals

Crystals now available to explore the physics

Our work at MIT: 1) The impurity question using x-rays 2) The spin excitations using neutrons Thursday, February 21, 13

H = J�

�ij⇥

⌅Si · ⌅Sj

Kagome antiferromagnet


H = J�

�ij⇥

⌅Si · ⌅Sj =

P. Fazekas and P. W. Anderson, Philos. Mag. 30, 23 (1974).



H = J�

�ij⇥

⌅Si · ⌅Sj =


P. Fazekas and P. W. Anderson, Philos. Mag. 30, 23 (1974).


Alternative view Pick a reference configuration



Alternative view A nearby configuration



Alternative view Difference: a closed loop



Alternative view Ground state: sum over closed loops



Mott insulator: Kagome antiferromagnet



| i ) Ground state of entire system,

⇢ = | ih |

⇢A = TrB⇢ = density matrix of region A

Entanglement entropy SE = �Tr (⇢A ln ⇢A)

B

Entanglement entropy

A P


| i ) Ground state of entire system,

⇢ = | ih |

Take | i = 1p2(|"iA |#iB � |#iA |"iB)

Then ⇢A = TrB⇢ = density matrix of region A=

12 (|"iA h"|A + |#iA h#|A)

Entanglement entropy SE = �Tr (⇢A ln ⇢A)= ln 2

Entanglement entropy


Band insulators

E

MetalMetal

carryinga current

InsulatorSuperconductor

kAn even number of electrons per unit cell

Entanglement entropy of a band insulator


SE = aP � b exp(�cP )

where P is the surface area (perimeter)

of the boundary between A and B.

B

Entanglement entropy of a band insulator

A P


B

Entanglement in the Z2 spin liquid

A

The sum over closed loops is characteristic of the Z2 spin liquid, introduced in N. Read and S. Sachdev, Phys. Rev. Lett. 66, 1773 (1991), X.-G. Wen, Phys. Rev. B 44, 2664 (1991)

P


B


A

Sum over closed loops: only an even number of links cross the boundary between A and B

The sum over closed loops is characteristic of the Z2 spin liquid, introduced in N. Read and S. Sachdev, Phys. Rev. Lett. 66, 1773 (1991), X.-G. Wen, Phys. Rev. B 44, 2664 (1991)

P


SE = aP � ln(2)

where P is the surface area (perimeter)

of the boundary between A and B.

B


A P

A. Hamma, R. Ionicioiu, and P. Zanardi, Phys. Rev. A 71, 022315 (2005) M. Levin and X.-G. Wen, Phys. Rev. Lett. 96, 110405 (2006); A. Kitaev and J. Preskill, Phys. Rev. Lett. 96, 110404 (2006)

Y. Zhang, T. Grover, and A. Vishwanath, Phys. Rev. B 84, 075128 (2011)Thursday, February 21, 13

Simeng Yan, D. A. Huse, and S. R. White, Science 332, 1173 (2011).

Mott insulator: Kagome antiferromagnet

S. Depenbrock, I. P. McCulloch, and U. Schollwoeck, Physical Review Letters 109, 067201 (2012)

Hong-Chen Jiang, Z. Wang, and L. Balents, Nature Physics 8, 902 (2012)

Strong numerical evidence for a Z2 spin liquid


Nature 492, 406 (2012)

LETTERdoi:10.1038/nature11659

Fractionalized excitations in the spin-liquid state of akagome-lattice antiferromagnetTian-Heng Han1, Joel S. Helton2, Shaoyan Chu3, Daniel G. Nocera4, Jose A. Rodriguez-Rivera2,5, Collin Broholm2,6 & Young S. Lee1

The experimental realization of quantum spin liquids is a long-sought goal in physics, as they represent new states of matter. Quan-tum spin liquids cannot be described by the broken symmetriesassociated with conventional ground states. In fact, the interactingmagnetic moments in these systems do not order, but are highlyentangled with one another over long ranges1. Spin liquids have aprominent role in theories describing high-transition-temperaturesuperconductors2,3, and the topological properties of these statesmay have applications in quantum information4. A key feature ofspin liquids is that they support exotic spin excitations carryingfractional quantum numbers. However, detailed measurements ofthese ‘fractionalized excitations’ have been lacking. Here we reportneutron scattering measurements on single-crystal samples of thespin-1/2 kagome-lattice antiferromagnet ZnCu3(OD)6Cl2 (also calledherbertsmithite), which provide striking evidence for this characte-ristic feature of spin liquids. At low temperatures, we find that thespin excitations form a continuum, in contrast to the conventionalspin waves expected in ordered antiferromagnets. The observation ofsuch a continuum is noteworthy because, so far, this signature offractional spin excitations has been observed only in one-dimensionalsystems. The results also serve as a hallmark of the quantum spin-liquid state in herbertsmithite.

In a spin liquid, the atomic magnetic moments are strongly corre-lated but do not order or freeze even in the limit as the temperature, T,goes to zero. Although many types of quantum spin-liquid states existin theory, a feature that is expected to be common to all is the presenceof deconfined spinons as an elementary excitation from the groundstate1. Spinons are spin-half (S 5 1/2) quantum excitations into whichconventional spin-wave excitations with S 5 1 fractionalize. In onedimension, this phenomenon is well established for the S 5 1/2Heisenberg antiferromagnetic chain, where spinons may be thoughtof as magnetic domain boundaries that disrupt Neel order and are freeto propagate away from each other. In the one-dimensional compoundKCuF3, a continuum of spinon excitations has been well characterizedusing neutron scattering5. In two dimensions, the nature of the spinonexcitations is less clear. First, the existence of two-dimensional magnetswith a quantum spin-liquid ground state is still a matter of great debate.Second, the various spin-liquid states which are proposed in theorygive rise to a variety of spinon excitation spectra, which may be eithergapped or gapless.

The S 5 1/2 kagome-lattice Heisenberg antiferromagnet has longbeen recognized as a promising system in which to search for quantumspin-liquid states, because the kagome network of corner-sharing tri-angles frustrates long-range magnetic order6–8. We have devised syn-thetic methods to produce herbertsmithite (ZnCu3(OH)6Cl2) in whichthe S 5 1/2 Cu21 moments are arranged on a structurally perfectkagome lattice9 and nonmagnetic Zn21 ions separate the lattice planes.A depiction of the crystal structure is shown in Supplementary Fig. 1.Whereas herbertsmithite typically contains a small percentage of excess

Cu21 ions (,5% of the total) substituting for Zn21 ions in the inter-layer sites, the kagome planes contain only Cu21 ions10. Measurementson powder samples11–13 indicate strong antiferromagnetic super-exchange (J < 17 meV, where J is the exchange coupling that appearsin the nearest-neighbour Heisenberg Hamiltonian) and the absenceof long-range magnetic order or spin freezing down to tempera-tures of T 5 0.05 K. The bulk magnetic properties reveal a smallDzyaloshinskii–Moriya interaction and an easy-axis exchangeanisotropy14,15, both of order J/10. Despite these small imperfections,the nearest-neighbour Heisenberg model on a kagome lattice is still anexcellent approximation of the spin Hamiltonian for herbertsmithite.This is especially important, because recent calculations on recordlattice sizes indicate that the ground state of this model is in fact aquantum spin liquid16. Thus, experiments to probe the spin correla-tions in herbertsmithite are all the more urgent.

To this end, we recently succeeded in developing a technique for thegrowth of large, high-quality single crystals of herbertsmithite17, andsmall pieces have been used in studies involving local probes18,19,anomalous X-ray diffraction10, susceptibility15 and Raman scattering20.In this Letter, we report inelastic neutron scattering measurements ona large, deuterated, single-crystal sample of herbertsmithite. The neu-tron scattering cross-section is directly proportional to the dynamicstructure factor Stot(Q, v) (where Q and v stand for the momentumand energy transferred to the sample, respectively), which includesboth the nuclear and magnetic signals. The magnetic part, Smag(Q, v),is the Fourier transform (in time and space) of the spin–spin correlationfunction and can be obtained by subtracting the nuclear scattering asdescribed in the Supplementary Information. After calibration withrespect to a vanadium standard, the measured structure factors areexpressed in absolute units.

Contour plots of Stot(Q, v) are shown in Fig. 1a–c for T 5 1.6 K andthree different energy transfers Bv (B denotes Planck’s constantdivided by 2p). Figure 1a shows data for Bv 5 6 meV. Surprisingly,the scattered intensity is exceedingly diffuse, spanning a large fractionof the hexagonal Brillouin zone. A similar pattern of diffuse scatter-ing is observed for Bv 5 2 meV (Fig. 1b). The diffuse nature of thescattering at a temperature that is two orders of magnitude below theexchange energy scale, J, is in strong contrast to observations in non-frustrated quantum magnets. The S 5 1/2 square-lattice antiferromag-net La2CuO4 develops substantial antiferromagnetic correlations forT , J/2 (ref. 21), temperatures at which the low-energy scattering isstrongly peaked near the (p,p) point in reciprocal space. In herbert-smithite, the scattered intensity is not strongly peaked at any spe-cific point, and this remains true for all energies measured fromBv 5 0.25 to 11 meV. This behaviour is also markedly differentfrom that observed in the larger, S 5 5/2 kagome antiferromagnetKFe3(OH)6(SO4)2 which becomes magnetically ordered at low tem-peratures and has magnetic peaks at q 5 0 wavevectors above theordering temperature22.

1Department of Physics, Massachusetts Institute of Technology, Cambridge, Massachusetts 02139, USA. 2NIST Center for Neutron Research, National Institute of Standards and Technology, Gaithersburg,Maryland 20899, USA. 3Center for Materials Science and Engineering, Massachusetts Institute of Technology, Cambridge, Massachusetts 02139, USA. 4Department of Chemistry, Massachusetts Instituteof Technology, Cambridge, Massachusetts 02139, USA. 5Department of Materials Science and Engineering, University of Maryland, College Park, Maryland 20742, USA. 6Institute for Quantum Matter andDepartment of Physics and Astronomy, The Johns Hopkins University, Baltimore, Maryland 21218, USA.

4 0 6 | N A T U R E | V O L 4 9 2 | 2 0 / 2 7 D E C E M B E R 2 0 1 2

Macmillan Publishers Limited. All rights reserved©2012







Outline


M. Greiner, O. Mandel, T. Esslinger, T. W. Hänsch, and I. Bloch, Nature 415, 39 (2002).

Ultracold 87Rbatoms - bosons

Superfluid-insulator transition


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0

! a complex field representing the

Bose-Einstein condensate of the superfluid


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0

system with a recently developed scheme based on single-atom-resolved detection24. It is the high sensitivity of this method thatallowed us to reduce the modulation amplitude by almost an orderof magnitude compared with earlier experiments20,21 and to stay wellwithin the linear response regime (Supplementary Information).

The results for selected lattice depths V0 are shown in Fig. 2b. Weobserve a gapped response with an asymmetric overall shape that willbe analysed in the following paragraphs. Notably, the maximumobserved temperature after modulation is well below the ‘melting’temperature for a Mott insulator in the atomic limit25, Tmelt < 0.2U/kB

(kB, Boltzmann’s constant), demonstrating that our experiments probethe quantum gas in the degenerate regime. To obtain numerical valuesfor the onset of spectral response, we fitted each spectrum with an errorfunction centred at a frequency n0 (Fig. 2b, black lines). With japproaching jc, the shift of the gap to lower frequencies is alreadyvisible in the raw data (Fig. 2b) and becomes even more apparent forthe fitted gap n0 as a function of j/jc (Fig. 2a, filled circles). The n0 valuesare in quantitative agreement with a prediction for the Higgs gap nSF atcommensurate filling (solid line):

hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2

Here h denotes Planck’s constant. This value is based on an analysis ofvariations around a mean-field state7,16 (throughout the manuscript,we have rescaled jc in the theoretical calculations to match the valuejc<0:06 obtained from quantum Monte Carlo simulations26).

The sharpness of the spectral onset can be quantified by the width ofthe fitted error function, which is shown as vertical dashed lines inFig. 2a. Approaching the critical point, the spectral onset becomessharper, and the width normalized to the centre frequency n0 remainsconstant (Supplementary Fig. 3). The constancy of this ratio indicatesthat the width of the spectral onset scales with the distance to thecritical point in the same way as the gap frequency.

We observe similar gapped responses in the Mott insulating regime(Supplementary Information and Fig. 5a), with the gap closing con-tinuously when approaching the critical point (Fig. 2a, open circles).We interpret this as a result of combined particle and hole excitationswith a frequency given by the Mott excitation gap that closes at thetransition point16. The fitted gaps are consistent with the Mott gap

hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2

where nMI is the Mott gap as predicted by mean-field theory16 (Fig. 2a,dashed line).

The observed softening of the onset of spectral response in thesuperfluid regime has led to an identification of the experimentalsignal with a response from collective excitations of Higgs type. Togain further insight into the full in-trap response, we calculated theeigenspectrum of the system in a Gutzwiller approach16,22 (Methodsand Supplementary Information). The result is a series of discreteeigenfrequencies (Fig. 3a), and the corresponding eigenmodes showin-trap superfluid density distributions, which are reminiscent of thevibrational modes of a drum (Fig. 3b). The frequency of the lowest-lying amplitude-like eigenmode n0,G closely follows the long-wave-length prediction for homogeneous commensurate filling nSF over awide range of couplings j/jc until the response rounds off in the vicinityof the critical point due to the finite size of the system (Fig. 3c). Fittingthe low-frequency edge of the experimental data can be interpreted asextracting the frequency of this mode, which explains the goodquantitative agreement with the prediction for the homogeneous com-mensurate filling in Fig. 2a. Modes at different frequencies from thelowest-lying amplitude-like mode broaden the spectrum only abovethe onset of spectral response.

An eigenmode analysis, however, does not yield any informationabout the finite spectral width of the modes, which stems from theinteraction between amplitude and phase excitations. We will considerthe question of the spectral width by analysing the low-, intermediate-and high-frequency parts of the response separately. We begin byexamining the low-frequency part of the response, which is expectedto be governed by a process coupling a virtually excited amplitudemode to a pair of phase modes with opposite momenta. As a result,the response of a strongly interacting, two-dimensional superfluid is

a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)

Lattice loading Modulation Hold time Ramp to atomic limit

Temperaturemeasurement

V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ

Figure 1 | Illustration of the Higgs mode and experimental sequence.a, Classical energy density V as a function of the order parameter Y. Within theordered (superfluid) phase, Nambu–Goldstone and Higgs modes arise fromphase and amplitude modulations (blue and red arrows in panel 1). As thecoupling j 5 J/U (see main text) approaches the critical value jc, the energydensity transforms into a function with a minimum at Y 5 0 (panels 2 and 3).Simultaneously, the curvature in the radial direction decreases, leading to acharacteristic reduction of the excitation frequency for the Higgs mode. In thedisordered (Mott insulating) phase, two gapped modes exist, respectivelycorresponding to particle and hole excitations in our case (red and blue arrow inpanel 3). b, The Higgs mode can be excited with a periodic modulation of thecoupling j, which amounts to a ‘shaking’ of the classical energy densitypotential. In the experimental sequence, this is realized by a modulation of theoptical lattice potential (see main text for details). t 5 1/nmod; Er, lattice recoilenergy.

LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Particles and holes correspond

to the 2 normal modes in the

oscillation of about = 0.

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


Insulator (the vacuum) at large repulsion between bosons


Particles ⇠ †

Excitations of the insulator:


Holes ⇠

Excitations of the insulator:


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0




hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Particles and holes correspond

to the 2 normal modes in the

oscillation of about = 0.

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0




hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Nambu-Goldstone mode is the

oscillation in the phase

at a constant non-zero | |.

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

A conformal field theory

in 2+1 spacetime dimensions:

a CFT3

h i 6= 0 h i = 0

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

Quantum state with

complex, many-body,

“long-range” quantum entanglement

h i 6= 0 h i = 0

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

No well-defined normal modes,

or particle-like excitations

h i 6= 0 h i = 0

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0




hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Higgs mode is theoscillation in theamplitude | |. This decaysrapidly by emitting pairsof Nambu-Goldstone modes.

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

h i 6= 0 h i = 0




hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Despite rapid decay,there is a well-defined

Higgs “quasi-normal mode”.This is associated with a pole

in the lower-halfof the complex frequency plane.

A. V. Chubukov, S. Sachdev, and J. Ye, PRB 49,11919 (1994). S. Sachdev, PRB 59, 14054 (1999).

W. Zwerger, PRL 92, 027203 (2004).

D. Podolsky, A. Auerbach, and D. P. Arovas, PRB 84, 174522 (2011).

D. Podolsky and S. Sachdev, PRB 86, 054508 (2012). L. Pollet and N. Prokof’ev, PRL 109, 010401 (2012).

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


The Higgs quasi-normal mode is at the frequency

!pole

�= �i

4

⇡+

1

N

16�4 +

p2 log

�3� 2

p2��

⇡2

+ 2.46531203396 i

!+O

✓1

N2

◆

where � is the particle gap at the complementary point in the “paramagnetic” statewith the same value of |�� c|, and N = 2 is the number of vector components of .The universal answer is a consequence of the strong interactions in the CFT3

!

D. Podolsky and S. Sachdev, Phys. Rev. B 86, 054508 (2012).

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


expected to diverge at low frequencies, if the probe in use coupleslongitudinally to the order parameter2,4,5,9 (for example to the real partof Y, if the equilibrium value of Y was chosen along the real axis), as isthe case for neutron scattering. If, instead, the coupling is rotationallyinvariant (for example through coupling to jYj2), as expected forlattice modulation, such a divergence could be avoided and the

response is expected to scale as n3 at low frequencies3,6,9,17.Combining this result with the scaling dimensions of the responsefunction for a rotationally symmetric perturbation coupling to jYj2,we expect the low-frequency response to be proportional to(1 2 j/jc)

22n3 (ref. 9 and Methods). The experimentally observed sig-nal is consistent with this scaling at the ‘base’ of the absorption feature(Fig. 4). This indicates that the low-frequency part is dominated byonly a few in-trap eigenmodes, which approximately show the genericscaling of the homogeneous system for a response function describingcoupling to jYj2.

In the intermediate-frequency regime, it remains a challenge toconstruct a first-principles analytical treatment of the in-trap systemincluding all relevant decay and coupling processes. Lacking such atheory, we constructed a heuristic model combining the discrete spec-trum from the Gutzwiller approach (Fig. 3a) with the line shape for ahomogeneous system based on an O(N) field theory in two dimen-sions, calculated in the large-N limit3,6 (Methods). An implicit assump-tion of this approach is a continuum of phase modes, which is

Sup

er!u

id d

ensi

ty

0.1

0.3

0.5

0.7

a

b1

0 0.5 1 1.5 2

0

0.02

0.04

d

0 0.5 0.6 0.7 0.8 0.9 1 1.1

Q0,G

0

0.5

1

Line

str

engt

h (a

.u.)

hQmod/U

hQ/U

j/jc

c

1 1.5 2 2.50

0.4

0.8

1.2V0 = 9.5Erj/jc = 1.4

k BΔT

/U, Δ

E/U

Q0,G

2

2

1

3

3

4

4

hQmod/U

Figure 3 | Theory of in-trap response. a, A diagonalization of the trappedsystem in a Gutzwiller approximation shows a discrete spectrum of amplitude-like eigenmodes. Shown on the vertical axis is the strength of the response to amodulation of j. Eigenmodes of phase type are not shown (Methods) and n0,G

denotes the gap as calculated in the Gutzwiller approximation. a.u., arbitraryunits. b, In-trap superfluid density distribution for the four amplitude modeswith the lowest frequencies, as labelled in a. In contrast to the superfluiddensity, the total density of the system stays almost constant (not shown).c, Discrete amplitude mode spectrum for various couplings j/jc. Each red circlecorresponds to a single eigenmode, with the intensity of the colour beingproportional to the line strength. The gap frequency of the lowest-lying modefollows the prediction for commensurate filling (solid line; same as in Fig. 2a)until a rounding off takes place close to the critical point due to the finite size ofthe system. d, Comparison of the experimental response at V0 5 9.5Er (bluecircles and connecting blue line; error bars, s.e.m.) with a 2 3 2 cluster mean-field simulation (grey line and shaded area) and a heuristic model (dashed line;for details see text and Methods). The simulation was done for V0 5 9.5Er (greyline) and for V0 5 (1 6 0.02) 3 9.5Er (shaded grey area), to account for theexperimental uncertainty in the lattice depth, and predicts the energyabsorption per particle DE.

0 0.2 0.4 0.6 0.8 1

0

0.01

0.02

0.03

Qmod/U

(1 –

j/j c)

2 kBΔT

/U

0 0.2 0.4 0.6 0.8 1

0

0.01

0.02

0.03

Qmod/U

k BΔT

/U

Figure 4 | Scaling of the low-frequency response. The low-frequencyresponse in the superfluid regime shows a scaling compatible with theprediction (1 2 j/jc)

22n3 (Methods). Shown is the temperature responserescaled with (1 2 j/jc)

2 for V0 5 10Er (grey), 9.5Er (black), 9Er (green), 8.5Er

(blue) and 8Er (red) as a function of the modulation frequency. The black line isa fit of the form anb with a fitted exponent b 5 2.9(5). The inset shows the samedata points without rescaling, for comparison. Error bars, s.e.m.

hQ0/U

j/jc

0 0.5 1 1.5 2 2.50

0.2

0.4

0.6

0.8

1

1.20 0.03 0.06 0.09 0.12 0.15

j

Super!uidMott Insulator

a b

3

1

2

V0 = 8Erj/jc = 2.2

k BT/U

1

2

3

V0 = 9Erj/jc = 1.6

V0 = 10Erj/jc = 1.2

Q0

0.11

0.13

0.15

0.17

0.12

0.14

0.16

0.18

0 400 8000.12

0.14

0.16

0.18

Qmod (Hz)

Q0

Q0

Figure 2 | Softening of the Higgs mode. a, The fitted gap values hn0/U(circles) show a characteristic softening close to the critical point in quantitativeagreement with analytic predictions for the Higgs and the Mott gap (solid lineand dashed line, respectively; see text). Horizontal and vertical error barsdenote the experimental uncertainty of the lattice depths and the fit error for thecentre frequency of the error function, respectively (Methods). Vertical dashedlines denote the widths of the fitted error function and characterize thesharpness of the spectral onset. The blue shading highlights the superfluid

region. b, Temperature response to lattice modulation (circles and connectingblue line) and fit with an error function (solid black line) for the three differentpoints labelled in a. As the coupling j approaches the critical value jc, the changein the gap values to lower frequencies is clearly visible (from panel 1 to panel 3).Vertical dashed lines mark the frequency U/h corresponding to the on-siteinteraction. Each data point results from an average of the temperatures over,50 experimental runs. Error bars, s.e.m.

RESEARCH LETTER

4 5 6 | N A T U R E | V O L 4 8 7 | 2 6 J U L Y 2 0 1 2





hnSF=U~ 3!!!2p

{4" #

1zj=jc! "$ %1=2

j=jc{1! "1=2




hnMI=U~ 1z 12!!!2p

{17" #

j=jc$ %1=2

1{j=jc! "1=2




a1

2

V

Re( )Im( )

Higgs modeNambu–

Goldstonemode

j/jc 1

0 100 200 300 4000

5

10

15

20

Time (ms)

Lat

tice

dept

h (E

r)



V0

Ttot = 200 ms

A = 0.03V0

Tmod = 20WW

b

3

j/jc * 1

j/jc , 1

ΨΨ


LETTER RESEARCH

2 6 J U L Y 2 0 1 2 | V O L 4 8 7 | N A T U R E | 4 5 5


Observation of Higgs quasi-normal mode across the superfluid-insulator transition of ultracold atoms in a 2-dimensional optical lattice:Response to modulation of lattice depth scales as expected from the LHP pole

Manuel Endres, Takeshi Fukuhara, David Pekker, Marc Cheneau, Peter Schaub, Christian Gross, Eugene Demler, Stefan Kuhr, and Immanuel Bloch, Nature 487, 454 (2012).


expected to diverge at low frequencies, if the probe in use coupleslongitudinally to the order parameter2,4,5,9 (for example to the real partof Y, if the equilibrium value of Y was chosen along the real axis), as isthe case for neutron scattering. If, instead, the coupling is rotationallyinvariant (for example through coupling to jYj2), as expected forlattice modulation, such a divergence could be avoided and the

response is expected to scale as n3 at low frequencies3,6,9,17.Combining this result with the scaling dimensions of the responsefunction for a rotationally symmetric perturbation coupling to jYj2,we expect the low-frequency response to be proportional to(1 2 j/jc)

22n3 (ref. 9 and Methods). The experimentally observed sig-nal is consistent with this scaling at the ‘base’ of the absorption feature(Fig. 4). This indicates that the low-frequency part is dominated byonly a few in-trap eigenmodes, which approximately show the genericscaling of the homogeneous system for a response function describingcoupling to jYj2.

In the intermediate-frequency regime, it remains a challenge toconstruct a first-principles analytical treatment of the in-trap systemincluding all relevant decay and coupling processes. Lacking such atheory, we constructed a heuristic model combining the discrete spec-trum from the Gutzwiller approach (Fig. 3a) with the line shape for ahomogeneous system based on an O(N) field theory in two dimen-sions, calculated in the large-N limit3,6 (Methods). An implicit assump-tion of this approach is a continuum of phase modes, which is

Sup

er!u

id d

ensi

ty

0.1

0.3

0.5

0.7

a

b1

0 0.5 1 1.5 2

0

0.02

0.04

d

0 0.5 0.6 0.7 0.8 0.9 1 1.1

Q0,G

0

0.5

1

Line

str

engt

h (a

.u.)

hQmod/U

hQ/U

j/jc

c

1 1.5 2 2.50

0.4

0.8

1.2V0 = 9.5Erj/jc = 1.4

k BΔT

/U, Δ

E/U

Q0,G

2

2

1

3

3

4

4

hQmod/U

Figure 3 | Theory of in-trap response. a, A diagonalization of the trappedsystem in a Gutzwiller approximation shows a discrete spectrum of amplitude-like eigenmodes. Shown on the vertical axis is the strength of the response to amodulation of j. Eigenmodes of phase type are not shown (Methods) and n0,G

denotes the gap as calculated in the Gutzwiller approximation. a.u., arbitraryunits. b, In-trap superfluid density distribution for the four amplitude modeswith the lowest frequencies, as labelled in a. In contrast to the superfluiddensity, the total density of the system stays almost constant (not shown).c, Discrete amplitude mode spectrum for various couplings j/jc. Each red circlecorresponds to a single eigenmode, with the intensity of the colour beingproportional to the line strength. The gap frequency of the lowest-lying modefollows the prediction for commensurate filling (solid line; same as in Fig. 2a)until a rounding off takes place close to the critical point due to the finite size ofthe system. d, Comparison of the experimental response at V0 5 9.5Er (bluecircles and connecting blue line; error bars, s.e.m.) with a 2 3 2 cluster mean-field simulation (grey line and shaded area) and a heuristic model (dashed line;for details see text and Methods). The simulation was done for V0 5 9.5Er (greyline) and for V0 5 (1 6 0.02) 3 9.5Er (shaded grey area), to account for theexperimental uncertainty in the lattice depth, and predicts the energyabsorption per particle DE.

0 0.2 0.4 0.6 0.8 1

0

0.01

0.02

0.03

Qmod/U

(1 –

j/j c)

2 kBΔT

/U

0 0.2 0.4 0.6 0.8 1

0

0.01

0.02

0.03

Qmod/U

k BΔT

/U

Figure 4 | Scaling of the low-frequency response. The low-frequencyresponse in the superfluid regime shows a scaling compatible with theprediction (1 2 j/jc)

22n3 (Methods). Shown is the temperature responserescaled with (1 2 j/jc)

2 for V0 5 10Er (grey), 9.5Er (black), 9Er (green), 8.5Er

(blue) and 8Er (red) as a function of the modulation frequency. The black line isa fit of the form anb with a fitted exponent b 5 2.9(5). The inset shows the samedata points without rescaling, for comparison. Error bars, s.e.m.

hQ0/U

j/jc

0 0.5 1 1.5 2 2.50

0.2

0.4

0.6

0.8

1

1.20 0.03 0.06 0.09 0.12 0.15

j

Super!uidMott Insulator

a b

3

1

2

V0 = 8Erj/jc = 2.2

k BT/U

1

2

3

V0 = 9Erj/jc = 1.6

V0 = 10Erj/jc = 1.2

Q0

0.11

0.13

0.15

0.17

0.12

0.14

0.16

0.18

0 400 8000.12

0.14

0.16

0.18

Qmod (Hz)

Q0

Q0

Figure 2 | Softening of the Higgs mode. a, The fitted gap values hn0/U(circles) show a characteristic softening close to the critical point in quantitativeagreement with analytic predictions for the Higgs and the Mott gap (solid lineand dashed line, respectively; see text). Horizontal and vertical error barsdenote the experimental uncertainty of the lattice depths and the fit error for thecentre frequency of the error function, respectively (Methods). Vertical dashedlines denote the widths of the fitted error function and characterize thesharpness of the spectral onset. The blue shading highlights the superfluid

region. b, Temperature response to lattice modulation (circles and connectingblue line) and fit with an error function (solid black line) for the three differentpoints labelled in a. As the coupling j approaches the critical value jc, the changein the gap values to lower frequencies is clearly visible (from panel 1 to panel 3).Vertical dashed lines mark the frequency U/h corresponding to the on-siteinteraction. Each data point results from an average of the temperatures over,50 experimental runs. Error bars, s.e.m.

RESEARCH LETTER

4 5 6 | N A T U R E | V O L 4 8 7 | 2 6 J U L Y 2 0 1 2


Manuel Endres, Takeshi Fukuhara, David Pekker, Marc Cheneau, Peter Schaub, Christian Gross, Eugene Demler, Stefan Kuhr, and Immanuel Bloch, Nature 487, 454 (2012).

Observation of Higgs quasi-normal mode across the superfluid-insulator transition of ultracold atoms in a 2-dimensional optical lattice:Response to modulation of lattice depth scales as expected from the LHP pole

D. Podolsky and S. Sachdev, Phy. Rev. B 86, 054508 (2012).


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

A conformal field theory

in 2+1 spacetime dimensions:

a CFT3

h i 6= 0 h i = 0

S =

Zd2rdt

⇥|@t |2 � c2|rr |2 � V ( )

⇤

V ( ) = (�� c)| |2 + u�| |2

�2


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

CFT3 at T>0


Quantum critical dynamics

S. Sachdev, Quantum Phase Transitions, Cambridge (1999).

Quantum “nearly perfect fluid”

with shortest possible local equilibration time, ⌧eq

⌧eq = C ~kBT

where C is a universal constant.

Response functions are characterized by poles in LHPwith ! ⇠ kBT/~.

These poles (quasi-normal modes) appear naturally inthe holographic theory.

(Analogs of Higgs quasi-normal mode.)


M.P.A. Fisher, G. Grinstein, and S.M. Girvin, Phys. Rev. Lett. 64, 587 (1990) K. Damle and S. Sachdev, Phys. Rev. B 56, 8714 (1997).

� =

Q2

h⇥ [Universal constant O(1) ]

(Q is the “charge” of one boson)

Quantum critical dynamics

Transport co-oe�cients not determined

by collision rate of quasiparticles, but by

fundamental constants of nature

Conductivity


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

CFT3 at T>0


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

CFT3 at T>0

Boltzmann theory of particles/holes/vortices

does not apply


g

T

gc

0

InsulatorSuperfluid

Quantumcritical

TKT

��c

CFT3 at T>0

Needed: Accurate theory of

quantum critical dynamics







Outline


J. McGreevy, arXiv0909.0518

xir

Renormalization group: ) Follow coupling

constants of quantum many body theory as a func-

tion of length scale r


Key idea: ) Implement r as an extra dimen-

sion, and map to a local theory in d + 2 spacetime

dimensions.

r xi

Renormalization group: ) Follow coupling

constants of quantum many body theory as a func-

tion of length scale r

J. McGreevy, arXiv0909.0518


r

Holography

xi

For a relativistic CFT in d spatial dimensions, the

metric in the holographic space is fixed by de-

manding the scale transformation (i = 1 . . . d)

xi ! ⇣xi , t ! ⇣t , ds ! ds


r

Holography

xi

This gives the unique metric

ds

2=

1

r

2

��dt

2+ dr

2+ dx

2i

�

This is the metric of anti-de Sitter space AdSd+2.


zr

AdS4R

2,1

Minkowski

CFT3

AdS/CFT correspondence

xi


zr

AdS4R

2,1

Minkowski

CFT3

A

xi

Holography and Entanglement


zr

AdS4R

2,1

Minkowski

CFT3

A

S. Ryu and T. Takayanagi, Phys. Rev. Lett. 96, 18160 (2006).

Associate entanglement entropy with an observer in the enclosed spacetime region, who cannot observe “outside” : i.e. the region is surrounded by an imaginary horizon.

xi



zr

AdS4R

2,1

Minkowski

CFT3

A


The entropy of this region is bounded by its surface area (Bekenstein-Hawking-’t Hooft-Susskind)

xi



zr

AdS4R

2,1

Minkowski

CFT3

AMinimal

surface area measures

entanglemententropy


xi



zr

AdS4R

2,1

Minkowski

CFT3

A

• Computation of minimal surface area yields

SE = aP � �,where � is a shape-dependent universal number.

xi

M. A. Metlitski, C. A. Fuertes, and S. Sachdev, Physical Review B 80, 115122 (2009); H. Casini, M. Huerta, and R. Myers, JHEP 1105:036, (2011); I. Klebanov, S. Pufu, and B. Safdi, arXiv:1105.4598



zrThis emergent spacetime is a solution of Einstein gravity

with a negative cosmological constant

AdS4R

2,1

Minkowski

CFT3

AdS/CFT correspondence

SE =

Zd

4x

p�g

1

22

✓R+

6

L

2

◆�

xi


AdS/CFT correspondence at non-zero temperatures

AdS4-Schwarzschild black-brane

There is a family of solutions of Einstein

gravity which describe non-zero

temperatures

r

S =

Zd

4x

p�g

1

22

✓R+

6

L

2

◆�

Rh




r

Black-brane (horizon) at

temperature of 2+1 dimensional quantum

critical system

ds

2 =

✓L

r

◆2 dr

2

f(r)� f(r)dt2 + dx

2 + dy

2

�

with f(r) = 1� (r/Rh)3

Rh


ds

2 =

✓L

r

◆2 dr

2

f(r)� f(r)dt2 + dx

2 + dy

2

�

with f(r) = 1� (r/Rh)3



r



critical system

RhA 2+1

dimensional

system at its

quantum

critical point:

kBT =

3~4⇡Rh

.


Friction of quantum criticality = waves

falling past the horizon



D. T. Son



critical system

RhA 2+1

dimensional

system at its

quantum

critical point:

kBT =

3~4⇡Rh

.

r


Quasi-normal modes of waves near horizon --

quasi-normal modes of quantum criticality (and Higgs)





critical system

RhA 2+1

dimensional

system at its

quantum

critical point:

kBT =

3~4⇡Rh

.

r


AdS4 theory of quantum criticality

R. C. Myers, S. Sachdev, and A. Singh, Physical Review D 83, 066017 (2011)

D. Chowdhury, S. Raju, S. Sachdev, A. Singh, and P. Strack, arXiv:1210.5247

Most general e↵ective holographic theory for lin-ear charge transport with 4 spatial derivatives:

Sbulk =1

g

2M

Zd

4x

pg

1

4FabF

ab + �L

2CabcdF

abF

cd

�

+

Zd

4x

pg

� 1

22

✓R+

6

L

2

◆�,

This action is characterized by 3 dimensionless pa-rameters, which can be linked to data of the CFT(OPE coe�cients).



h

Q2�

0.0 0.5 1.0 1.5 2.00.0

0.5

1.0

1.5

⇥

4�T

� = 0

� =1

12

� = � 1

12


�(!)

�(1)


h

Q2�

0.0 0.5 1.0 1.5 2.00.0

0.5

1.0

1.5

⇥

4�T

� = 0

� =1

12

� = � 1

12


�(!)

�(1)


• Stability constraints on the e↵ective theory (|�| < 1/12)allow only a limited !-dependence in the conductiv-

ity. This contrasts with the Boltzmann theory in which

�(!)/�1 becomes very large in the regime of its validity.



Universal Scaling of the Conductivity at the Superfluid-Insulator Phase Transition

Jurij Smakov and Erik SørensenDepartment of Physics and Astronomy, McMaster University, Hamilton, Ontario L8S 4M1, Canada

(Received 30 May 2005; published 27 October 2005)

The scaling of the conductivity at the superfluid-insulator quantum phase transition in two dimensionsis studied by numerical simulations of the Bose-Hubbard model. In contrast to previous studies, we focuson properties of this model in the experimentally relevant thermodynamic limit at finite temperature T. Wefind clear evidence for deviations from !k scaling of the conductivity towards !k=T scaling at lowMatsubara frequencies !k. By careful analytic continuation using Pade approximants we show that thisbehavior carries over to the real frequency axis where the conductivity scales with !=T at smallfrequencies and low temperatures. We estimate the universal dc conductivity to be !! " 0:45#5$Q2=h,distinct from previous estimates in the T " 0, !=T % 1 limit.

DOI: 10.1103/PhysRevLett.95.180603 PACS numbers: 05.60.Gg, 02.70.Ss, 05.70.Jk

The nontrivial properties of materials in the vicinity ofquantum phase transitions [1] (QPTs) are an object ofintense theoretical [1–3] and experimental studies. Theeffect of quantum fluctuations driving the QPTs is espe-cially pronounced in low-dimensional systems, such ashigh-temperature superconductors and two-dimensional(2D) electron gases, exhibiting the quantum Hall effect.Particularly valuable are theoretical predictions of thebehavior of the dynamical response functions, such as theoptical conductivity and the dynamic structure factor, sincethey allow for direct comparison of the theoretical resultswith experimental data. It was pointed out by Damle andSachdev [2] that at the quantum-critical coupling thescaled dynamic conductivity T#2&d$=z!#!; T$ at low fre-quencies and temperatures is a function of the singlevariable @!=kBT:

!#!=T; T ! 0$ " #kBT=@c$#d&2$=z!Q!#@!=kBT$: (1)

Here !Q " Q2=h is the conductivity ‘‘quantum’’ (Q " 2efor the models we consider), !#x ' @!=kBT$ is a universaldimensionless scaling function, c a nonuniversal constant,and z the dynamical critical exponent. For d " 2 the ex-ponent vanishes, leading to a purely universal conductivity[4], depending only on frequency !, measured against acharacteristic time @" set by finite temperature T as@!=kBT. Once @!=kBT % 1, for fixed T, the system nolonger ‘‘feels’’ the effect of finite temperature and it isnatural to expect that at such high ! a crossover to atemperature-independent regime will take place [3], sothat !#!; T$ ( !#!$ with !#!$ decaying at high frequen-cies as 1=!2 [2]. Deviations from scaling of ! with !therefore signal that temperature effects have become im-portant. Note that the predicted universal behavior occursfor fixed !=T as T ! 0. The physical mechanisms oftransport are predicted [2] to be quite distinct in the differ-ent regimes determined by the value of the scaling variablex: hydrodynamic, collision dominated for x) 1, and col-lisionless, phase coherent for x% 1 with ! " !#1$

largely independent of x in d " 2 and ! independent ofT [2,5].

Intriguingly, early numerical studies [6–9] of QPTs inmodel systems have failed to observe scaling with@!=kBT. The results of the experiments seeking to verifythe scaling hypothesis are ambiguous as well. Some ofthem, performed at the 2D quantum Hall transitions [10]and 3D metal-insulator transitions [11], appear to supportit. Others either note the absence of scaling [12] or suggesta different scaling form [13]. While the discrepancy be-tween theory and experiment may be attributed to theunsuitable choice of the measurement regime [2], typicallyleading to @!=kBT % 1, there is no good reason why thepredicted scaling would not be observable in numericalsimulations if careful extrapolations first to L! 1 andthen T ! 0 for fixed !=T are performed.

Our primary goal is to resolve this controversy by per-forming precise numerical simulations of the frequency-dependent conductivity at finite temperatures in the vicin-ity of the 2D QPT, exploiting recent algorithmic advancesto access larger system sizes and a wider temperaturerange. After the extrapolation of the results to the thermo-dynamic and T " 0 limits and careful analytic continu-ation, we are able to demonstrate how the predicteduniversal behavior of the conductivity may indeed berevealed.

We consider the 2D Bose-Hubbard (BH) model with theHamiltonian H BH "H 0 *H 1, where the first termdescribes the noninteracting soft core bosons hopping viathe nearest-neighbor links of a 2D square lattice, and thesecond one includes the Hubbard-like on-site interactions:

H 0 " &tX

r;##byr br*# * byr*#br$ &$

Xrnr; (2)

H 1 "U2

Xrnr#nr & 1$: (3)

Here # " x; y, nr " byr br is the particle number operatoron site r, and byr ; br are the boson creation and annihilation

PRL 95, 180603 (2005) P H Y S I C A L R E V I E W L E T T E R S week ending28 OCTOBER 2005

0031-9007=05=95(18)=180603(4)$23.00 180603-1 ! 2005 The American Physical Society

The SSE data also display a high narrow peak at very lowfrequencies, whose position and shape are unstable withrespect to the choice of the initial image and MaxEntparameters. This is clearly an artifact of the method; how-ever, its presence is indicative of the tendency to accumu-late the weight at very low frequencies, in qualitativeagreement with H V result. The subsequent falloff in theconductivity at high frequencies is physically consistent,but its functional form depends on the Pade approximantused. For !=!c * 1=2, we expect the analytic continu-ation of the data for H V to become sensitive to the order ofthe approximant used and we therefore indicate the resultsin this regime by dotted lines only. We note that results atall temperatures yield the same dc conductivity !? !0:45"5#!Q, theoretically predicted [4] to be universal.Because of the very different scaling procedure this resultdiffers from previous numerical result !? ! 0:285"20#!Qon the same model [6] in the T ! 0 limit. It also differssignificantly from a theoretical estimate [2], !$ !1:037!Q, valid to leading order in " ! 3% d. Remark-ably, our result for the dc conductivity is very close to theone obtained in Ref. [8] for the phase transition in thedisordered Bose-Hubbard model. Experimental results in-dicate a value close to unity [26]; however, it was previ-ously observed [7] that long-range Coulomb interactions,impossible to include in the present study, tend to increase! considerably. The same data are shown versus !=T inFig. 4(b). Notably, when using this parametrization !ccancels out and all our data follow the same functionalform. The scaling with !=T at low frequencies is nowimmediately apparent, with a surprisingly wide low !=Tpeak. The width of this peak is consistent with the data inFig. 3(d). Furthermore, on the same !=T scale the con-tinuous time SSE data for H BH and the results for H Vqualitatively agree.

In summary, we have demonstrated that by doing a verycareful data analysis it is possible to observe the theoreti-cally predicted universal !=T scaling at the 2D superfluid-

insulator transition. We have also estimated the universaldc conductivity at this transition and found that it differssignificantly from existing numerical and theoreticalestimates.

We thank S. Sachdev, S. Girvin, and A. P. Young forvaluable comments and critical remarks. Financial sup-port from SHARCNET, NSERC, and CFI is gratefullyacknowledged. All calculations were done at theSHARCNET facility at McMaster University.

[1] J. Hertz, Phys. Rev. B 14, 1165 (1976); A. J. Millis, Phys.Rev. B 48, 7183 (1993); S. Sachdev, Quantum PhaseTransitions (Cambridge University Press, Cambridge,England, 2001); M. Vojta, Rep. Prog. Phys. 66, 2069(2003).

[2] K. Damle and S. Sachdev, Phys. Rev. B 56, 8714 (1997).[3] S. L. Sondhi et al., Rev. Mod. Phys. 69, 315 (1997).[4] M. P. A. Fisher, G. Grinstein, and S. M. Girvin, Phys. Rev.

Lett. 64, 587 (1990).[5] S. Sachdev, Phys. Rev. B 55, 142 (1997).[6] M.-C. Cha et al., Phys. Rev. B 44, 6883 (1991).[7] E. S. Sørensen et al., Phys. Rev. Lett. 69, 828 (1992);

M. Wallin et al., Phys. Rev. B 49, 12 115 (1994).[8] G. G. Batrouni et al., Phys. Rev. B 48, 9628 (1993).[9] R. T. Scalettar, N. Trivedi, and C. Huscroft, Phys. Rev. B

59, 4364 (1999).[10] L. W. Engel et al., Phys. Rev. Lett. 71, 2638 (1993).[11] H.-K. Lee et al., Phys. Rev. Lett. 80, 4261 (1998); Science

287, 633 (2000).[12] N. Q. Balaban, U. Meirav, and I. Bar-Joseph, Phys. Rev.

Lett. 81, 4967 (1998).[13] Y. S. Lee et al., Phys. Rev. B 66, 041104(R) (2002).[14] M. P. A. Fisher et al., Phys. Rev. B 40, 546 (1989).[15] A. W. Sandvik, J. Phys. A 25, 3667 (1992).[16] O. F. Syljuasen and A. W. Sandvik, Phys. Rev. E 66,

046701 (2002); O. F. Syljuasen, ibid. 67, 046701 (2003).[17] M. P. A. Fisher and D. H. Lee, Phys. Rev. B 39, 2756

(1989).[18] F. Alet and E. S. Sørensen, Phys. Rev. E 67, 015701(R)

(2003); 68, 026702 (2003).[19] W. Krauth and N. Trivedi, Europhys. Lett. 14, 627 (1991).[20] J. K. Freericks and H. Monien, Phys. Rev. B 53, 2691

(1996); N. Elstner and H. Monien, ibid. 59, 12 184 (1999).[21] G. D. Mahan, Many-Particle Physics (Plenum, New York,

1990), 2nd ed..[22] D. J. Scalapino, S. R. White, and S. Zhang, Phys. Rev.

Lett. 68, 2830 (1992); Phys. Rev. B 47, 7995 (1993).[23] R. Fazio and D. Zappala, Phys. Rev. B 53, R8883 (1996).[24] M. Jarrell and J. E. Gubernatis, Phys. Rep. 269, 133

(1996).[25] W. H. Press et al., Numerical Recipes in C (Cambridge

University Press, Cambridge, England, 1992).[26] See Y. Liu et al., Physica (Amsterdam) 83D, 163 (1995),

and references therein.[27] Finite Size Scaling and Numerical Simulations of Statis-

tical Systems, edited by V. Privman (World Scientific,Singapore, 1990).

0

0.2

0.4

0.6

0.8

!(’"

)

0 0.5 1 1.5"/"c

L#=32L#=40L#=48L#=64SSE

0

0.2

0.4

0.6

0.8

!(’"

)T/

0 25 50 75"/T

L#=32L#=40L#=48L#=64SSE

(a) (b)

FIG. 4. The real part of the conductivity !0 at the criticalcoupling in units of !Q. The data marked L#, plotted vs!=!c, were obtained using H V , combined with the analyticcontinuation of $"!=!c# as explained in the text. Results for!=!c * 1=2 are denoted by dotted lines. The data marked SSE,plotted vs !=10, were obtained by direct SSE simulations ofH BH with L ! 20, % ! 10 and subsequent maximum entropyanalysis (a). Results as a function of !=T (b).


180603-4

!!i!k"=!2"!Q" #h$kxi$!xx!i!k"

!k% #!i!k"

!k: (7)

Here h$kxi is the kinetic energy per link and #!i!k" is thefrequency-dependent stiffness. To measure !xx!i!k" wenote that !xx!$" may be expressed in terms of the correla-tion functions !%&

xx !r; $" # hK%x !r; $"K&

x !0; 0"i of operatorsK&x !r; $" # tbyr&x!$"br!$" and K$x !r; $" # tbyr !$"br&x!$",which may be estimated efficiently in SSE [15].Remarkably, it is possible to analytically perform theFourier transform with respect to $ yielding

!%&xx !r; !k" #

!1

'

Xn$2

m#0

"amn!!k"N!&;%;m""; (8)

where N!&;%;m" is the number of times the operatorsK%!r" and K&!0" appear in the SSE operator sequenceseparated by m operator positions, and n is the expansionorder. The coefficients "amn!!k" are given by the degeneratehypergeometric (Kummer) function: "amn!!k"#1 F1!m&1; n;$i'!k". This expression and (8) allow us to directlyevaluate !xx!r; !k" as a function of Matsubara frequencies,eliminating any errors associated with the discretization ofthe imaginary time interval. Analogously, in the link-current representation #!i!k" can be calculated [7], andthe conductivity can be obtained from Eq. (7).

In Fig. 3 we show results for !!i!k" versus !k obtainedusing the geometrical worm algorithm on H V at Kc[Fig. 3(a)] and by SSE simulations at tc;(c of H BH[Fig. 3(c)]. In both cases the results have been extrapolatedto the thermodynamic limit L! 1 at fixed '. As evidentfrom Fig. 3(a), the results deviate from scaling with !k atsmall !k and more significantly so at higher temperatures(small L$). These deviations are also visible in the con-tinuous time SSE data in Fig. 3(c), demonstrating that theycannot be attributed to time discretization errors. Similardeviations have been noted previously [6,7] but were notanalyzed at fixed '. Since the deviations persist in the L!1 limit at fixed ', they may only be interpreted as finite Teffects. Expecting a crossover to!k=T scaling at small!k,we plot our results versus !k=T in Fig. 3(d). For L$ ' 32,!!!1=T" is already independent of T (L$). In fact, asshown in Fig. 3(d), for !1...5, !!!k=T; T" can unambigu-ously be extrapolated to a finite !!!k=T; T ! 0" ( #!x"limit. This fact is a clear indication that !k=T scalingindeed occurs as T ! 0. Tentatively, for increasing!k=T, !!!k=T; T ! 0" appears to reach a constant valueof roughly 0:33!Q (#!1" in excellent agreement withtheoretical estimates [2,23]. We note that deviations from!k scaling appear to be largely absent in simulations ofH BH with disorder [7,8]. However, at this QCP the dy-namical critical exponent is different (z # 2). As is evidentfrom the size of the error bars in Fig. 3, simulations of H Vare much more efficient than the SSE simulations directlyon H BH. In the following analytic continuation we there-fore use the SSE data mostly as a consistency check.

Our results on the imaginary frequency axis are limitedby the lowest Matsubara frequency, !1 # 2"[email protected], the information about the behavior of !0!!" %Re!!!" at low! is embedded in values of the CCCF at allMatsubara frequencies, allowing us to determine it. Inorder to study the !=T scaling predicted for the hydro-dynamic collision-dominated regime [2] @!=kBT ) 1, wehave attempted analytic continuations of #!i!k" to obtain!0!!" at real frequencies. SSE results for H BH wereanalytically continued using the Bryan maximum entropy(ME) method [24] with flat initial image. For the resultsobtained for the link-current model H V we use a methodthat should be most sensitive to low frequencies !=!c < 1or #!x) 1". We fit the extrapolated low frequency part(first 10–15 Matsubara frequencies) of #!i!k" to a 6th-order polynomial. The resulting 6 coefficients are thenused to obtain a !3; 3" Pade approximant using standardtechniques [25]. This approximant is then used for theanalytic continuation of # by i!k ! !& i). Resultingreal frequency conductivities !0!!" are displayed inFig. 4(a) versus !=!c. The typical SSE data are plottedversus !=10 and are only shown for L # 20, ' # 10.

The results for H V show a broadened peak as !! 0,due to inelastic scattering, followed by a second peaknicely consistent in height and width with the SSE data.

0.2

0.3

0.4

!(i"

k)

0 0.1 0.2 0.3 0.4 0.5"k/(2#"c)

L$=16

L$=24

L$=32

L$=40

L$=48

L$=64

0 25 50 75 100"k/T

L$=16

L$=24

L$=32

L$=40

L$=48

T=0 Ext.

0.1

0.2

0.3

0.4!(

i"k)

0 10 20 30"k

%=2%=4%=6%=8%=10

0.006

0.007

0.008

0.009

0.010

&("

1)

100 200 300L

L$=64

Fit

(a)(b)

(c) (d)

FIG. 3 (color online). The conductivity !!!k" in units of !Qvs Matsubara frequency !k=!c as obtained from H V (a). Allresults have been extrapolated to the thermodynamic limit L!1 using the scaling form f!L" # a& b exp!$L=*"=

####Lp

[27] bycalculating #!!k" at fixed L$ using 9 lattice sizes from L #L$ . . . 4L$ as shown in (b). !!!k" in units of !Q vs Matsubarafrequency !k as obtained from SSE calculations of H BH, withsome typical error bars shown. All results have been extrapolatedto the thermodynamic limit by calculating !xx!!k" for fixed 'using 5 lattice sizes L # 12; . . . ; 30 (c). Scaling plot of theconductivity data from (a) vs !k=T. ! denotes extrapola-tions to T ! 0 (L$ ! 1) at fixed !k=T using: f!L$" # c&d exp!$L$=*$"=

######L$p

[27] (d).


180603-3



Universal Scaling of the Conductivity at the Superfluid-Insulator Phase Transition

Jurij Smakov and Erik SørensenDepartment of Physics and Astronomy, McMaster University, Hamilton, Ontario L8S 4M1, Canada

(Received 30 May 2005; published 27 October 2005)

The scaling of the conductivity at the superfluid-insulator quantum phase transition in two dimensionsis studied by numerical simulations of the Bose-Hubbard model. In contrast to previous studies, we focuson properties of this model in the experimentally relevant thermodynamic limit at finite temperature T. Wefind clear evidence for deviations from !k scaling of the conductivity towards !k=T scaling at lowMatsubara frequencies !k. By careful analytic continuation using Pade approximants we show that thisbehavior carries over to the real frequency axis where the conductivity scales with !=T at smallfrequencies and low temperatures. We estimate the universal dc conductivity to be !! " 0:45#5$Q2=h,distinct from previous estimates in the T " 0, !=T % 1 limit.

DOI: 10.1103/PhysRevLett.95.180603 PACS numbers: 05.60.Gg, 02.70.Ss, 05.70.Jk

The nontrivial properties of materials in the vicinity ofquantum phase transitions [1] (QPTs) are an object ofintense theoretical [1–3] and experimental studies. Theeffect of quantum fluctuations driving the QPTs is espe-cially pronounced in low-dimensional systems, such ashigh-temperature superconductors and two-dimensional(2D) electron gases, exhibiting the quantum Hall effect.Particularly valuable are theoretical predictions of thebehavior of the dynamical response functions, such as theoptical conductivity and the dynamic structure factor, sincethey allow for direct comparison of the theoretical resultswith experimental data. It was pointed out by Damle andSachdev [2] that at the quantum-critical coupling thescaled dynamic conductivity T#2&d$=z!#!; T$ at low fre-quencies and temperatures is a function of the singlevariable @!=kBT:

!#!=T; T ! 0$ " #kBT=@c$#d&2$=z!Q!#@!=kBT$: (1)

Here !Q " Q2=h is the conductivity ‘‘quantum’’ (Q " 2efor the models we consider), !#x ' @!=kBT$ is a universaldimensionless scaling function, c a nonuniversal constant,and z the dynamical critical exponent. For d " 2 the ex-ponent vanishes, leading to a purely universal conductivity[4], depending only on frequency !, measured against acharacteristic time @" set by finite temperature T as@!=kBT. Once @!=kBT % 1, for fixed T, the system nolonger ‘‘feels’’ the effect of finite temperature and it isnatural to expect that at such high ! a crossover to atemperature-independent regime will take place [3], sothat !#!; T$ ( !#!$ with !#!$ decaying at high frequen-cies as 1=!2 [2]. Deviations from scaling of ! with !therefore signal that temperature effects have become im-portant. Note that the predicted universal behavior occursfor fixed !=T as T ! 0. The physical mechanisms oftransport are predicted [2] to be quite distinct in the differ-ent regimes determined by the value of the scaling variablex: hydrodynamic, collision dominated for x) 1, and col-lisionless, phase coherent for x% 1 with ! " !#1$

largely independent of x in d " 2 and ! independent ofT [2,5].

Intriguingly, early numerical studies [6–9] of QPTs inmodel systems have failed to observe scaling with@!=kBT. The results of the experiments seeking to verifythe scaling hypothesis are ambiguous as well. Some ofthem, performed at the 2D quantum Hall transitions [10]and 3D metal-insulator transitions [11], appear to supportit. Others either note the absence of scaling [12] or suggesta different scaling form [13]. While the discrepancy be-tween theory and experiment may be attributed to theunsuitable choice of the measurement regime [2], typicallyleading to @!=kBT % 1, there is no good reason why thepredicted scaling would not be observable in numericalsimulations if careful extrapolations first to L! 1 andthen T ! 0 for fixed !=T are performed.

Our primary goal is to resolve this controversy by per-forming precise numerical simulations of the frequency-dependent conductivity at finite temperatures in the vicin-ity of the 2D QPT, exploiting recent algorithmic advancesto access larger system sizes and a wider temperaturerange. After the extrapolation of the results to the thermo-dynamic and T " 0 limits and careful analytic continu-ation, we are able to demonstrate how the predicteduniversal behavior of the conductivity may indeed berevealed.

We consider the 2D Bose-Hubbard (BH) model with theHamiltonian H BH "H 0 *H 1, where the first termdescribes the noninteracting soft core bosons hopping viathe nearest-neighbor links of a 2D square lattice, and thesecond one includes the Hubbard-like on-site interactions:

H 0 " &tX

r;##byr br*# * byr*#br$ &$

Xrnr; (2)

H 1 "U2

Xrnr#nr & 1$: (3)

Here # " x; y, nr " byr br is the particle number operatoron site r, and byr ; br are the boson creation and annihilation


0031-9007=05=95(18)=180603(4)$23.00 180603-1 ! 2005 The American Physical Society

QMC yields �(0)/�1 ⇡ 1.36

Holography yields �(0)/�1 = 1 + 4� with |�| 1/12.

Maximum possible holographic value �(0)/�1 = 1.33

W. Witzack-Krempa and S. Sachdev, arXiv:1302.0847Thursday, February 21, 13

Traditional CMT

Identify quasiparticles and their dispersions

Compute scattering matrix elements of quasiparticles (or of collective modes)

These parameters are input into a quantum Boltzmann equation

Deduce dissipative and dynamic properties at non-zero temperatures


Traditional CMT

Identify quasiparticles and their dispersions

Compute scattering matrix elements of quasiparticles (or of collective modes)

These parameters are input into a quantum Boltzmann equation

Deduce dissipative and dynamic properties at non-zero temperatures

Start with strongly interacting CFT without particle- or wave-like excitations

Compute OPE co-efficients of operators of the CFT

Relate OPE co-efficients to couplings of an effective graviational theory on AdS

Solve Einsten-Maxwell equations. Dynamics of quasi-normal modes of black branes.

Holography and black-branes







Outline


BaFe2(As1�x

Px

)2

K. Hashimoto, K. Cho, T. Shibauchi, S. Kasahara, Y. Mizukami, R. Katsumata, Y. Tsuruhara, T. Terashima, H. Ikeda, M. A. Tanatar, H. Kitano, N. Salovich, R. W. Giannetta, P. Walmsley, A. Carrington, R. Prozorov, and Y. Matsuda, Science 336, 1554 (2012).

Resistivity⇠ ⇢0 +ATn

FL


BaFe2(As1�x

Px

)2

K. Hashimoto, K. Cho, T. Shibauchi, S. Kasahara, Y. Mizukami, R. Katsumata, Y. Tsuruhara, T. Terashima, H. Ikeda, M. A. Tanatar, H. Kitano, N. Salovich, R. W. Giannetta, P. Walmsley, A. Carrington, R. Prozorov, and Y. Matsuda, Science 336, 1554 (2012).

Resistivity⇠ ⇢0 +ATn

FL

StrangeMetal

Metal


Electrons (fermions) occupy states inside a

Fermi “surface” (circle) of radius kF which is

determined by the density of electrons, Q.

The Metal

kx

ky �� kF !


Can bosons form a metal ?

Yes, if each boson, b, fractionalizes into 2 fermions (‘quarks’)

b = f1f2 !

• Each quark is charged under an emergent gauge force,

which encapsulates the entanglement in the ground state.

• The quarks have “hidden” Fermi surfaces of radius kF .

A Strange Metal




b = f1f2 !




O. I. Motrunich and M. P. A. Fisher, Phys. Rev. B 75, 235116 (2007)L. Huijse and S. Sachdev, Phys. Rev. D 84, 026001 (2011)

S. Sachdev, arXiv:1209.1637A Strange Metal




b = f1f2 !




kx

ky �� kF !

A Strange Metal


Holographic theory of a strange metal

Electric flux

The density of particles Q creates an electric flux Erwhich modifies the metric of the emergent spacetime.



The density of particles Q creates an electric flux Erwhich modifies the metric of the emergent spacetime.

Hidden Fermi surfaces

of “quarks” ?

Electric flux



Electric flux

The general metric transforms under rescaling as

xi ! ⇣ xi, t ! ⇣

zt, ds ! ⇣

✓/dds.

Recall: conformal matter has ✓ = 0, z = 1, and the metric is

anti-de Sitter


of “quarks” ?



Electric flux


xi ! ⇣ xi, t ! ⇣

zt, ds ! ⇣

✓/dds.


anti-de Sitter

The value ✓ = d�1 reproduces all the essential characteristicsof the entropy and entanglement entropy of a strange metal.

L. Huijse, S. Sachdev, B. Swingle, Physical Review B 85, 035121 (2012)

N. Ogawa, T. Takayanagi, and T. Ugajin, JHEP 1201,

125 (2012).


of “quarks” ?



Electric flux


xi ! ⇣ xi, t ! ⇣

zt, ds ! ⇣

✓/dds.


anti-de Sitter

L. Huijse, S. Sachdev, B. Swingle, Physical Review B 85, 035121 (2012)

N. Ogawa, T. Takayanagi, and T. Ugajin, JHEP 1201,

125 (2012).

The null-energy condition of gravity yields z � 1 + ✓/d. In d = 2, this

leads to z � 3/2. Field theory on strange metal yields z = 3/2 to 3 loops!

M. A. Metlitski and S. Sachdev, Phys. Rev. B 82, 075127 (2010)


of “quarks” ?


Conclusions

Realizations of many-particle entanglement:

Z2 spin liquids and conformal quantum critical points


New insights and solvable models for diffusion and transport of strongly interacting systems near quantum critical points

The description is far removed from, and complementary to, that of the quantum Boltzmann equation which builds on the quasiparticle/vortex picture.

Good prospects for experimental tests of frequency-dependent, non-linear, and non-equilibrium transport

Conclusions

Conformal quantum matter


Conclusions

More complex examples in metallic states are experimentally

ubiquitous, but pose difficult strong-coupling problems to conventional methods of field

theory


Conclusions

String theory and gravity in emergent dimensions

offer a remarkable new approach to describing states with many-particle quantum entanglement.

Much recent progress offers hope of a holographic description of “strange metals”


the quantum phases of matter and gauge-gravity dualityqpt.physics.harvard.edu/talks/ufl13.pdfthe...

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