On Holographic p-wave Superfluids with Back-reaction

On Holographic p-wave Superfluids with Back-reaction

Martin Ammon ammon@mppmu.mpg.de Johanna Erdmenger jke@mppmu.mpg.de Viviane Grass grass@mppmu.mpg.de Patrick Kerner pkerner@mppmu.mpg.de Andy O’Bannon ahob@mppmu.mpg.de Max-Planck-Institut für Physik (Werner-Heisenberg-Institut)
Föhringer Ring 6, 80805 München, Germany
July 15, 2019
Abstract

We numerically construct asymptotically Anti-de Sitter charged black hole solutions of -dimensional Einstein-Yang-Mills theory that, for sufficiently low temperature, develop vector hair. Via gauge-gravity duality, these solutions describe a strongly-coupled conformal field theory at finite temperature and density that undergoes a phase transition to a superfluid state with spontaneously broken rotational symmetry (a p-wave superfluid state). The bulk theory has a single free parameter, the ratio of the five-dimensional gravitational constant to the Yang-Mills coupling, which we denote as . Previous analyses have shown that in the so-called probe limit, where goes to zero and hence the gauge fields are ignored in Einstein’s equation, the transition to the superfluid state is second order. We construct fully back-reacted solutions, where is finite and the gauge fields are included in Einstein’s equation, and find that for values of above a critical value in units of the AdS radius, the transition becomes first order.

keywords:
Gauge/gravity duality, Black Holes, Phase transitions
Pacs:
11.25.Tq, 04.70.Bw, 05.70.Fh

1 Introduction

The Anti-de Sitter/Conformal Field Theory correspondence (AdS/CFT) Maldacena:1997re () provides a novel method for studying strongly-coupled systems at finite density. As such, it may have useful applications in condensed matter physics, especially for studying scale-invariant strongly-coupled systems, for example, low-temperature systems near quantum criticality (see for example refs.Herzog:2009xv (); Hartnoll:2009sz (); McGreevy:2009xe () and references therein). Such systems are not purely theoretical: the (thermo)dynamics of some high- superconductors may be controlled by a quantum critical point.

AdS/CFT is a holographic duality: it equates a weakly-coupled theory of gravity in -dimensional AdS space with a strongly-coupled -dimensional CFT “living” at the AdS boundary. CFT states with finite temperature are dual to black hole geometries, where the Hawking temperature of the black hole is identified with the temperature in the CFT Witten:1998zw (). The current of a global symmetry in the CFT will be dual to a gauge field in AdS space. AdS/CFT thus allows us to compute observables in a strongly-coupled CFT, in states with finite temperature and density, by studying asymptotically AdS charged black holes. By now AdS/CFT can model many basic phenomena in condensed matter physics, such as the quantum Hall effect Davis:2008nv (), non-relativistic scale-invariance Son:2008ye (); Balasubramanian:2008dm (), and Fermi surfaces Rey:2008zz (); Liu:2009dm (); Cubrovic:2009ye ().

AdS/CFT can also describe phase transitions to superfluid states, that is, phase transitions in which a sufficiently large charge density triggers spontaneous breaking of the symmetry: an operator charged under the acquires a nonzero expectation value Gubser:2008px (); Hartnoll:2008vx (); Hartnoll:2008kx (). We will refer to this as the operator “condensing.” The simplest bulk action that can describe such a transition is Einstein-Maxwell theory coupled to a charged scalar. In the bulk, a charged black hole develops scalar hair. In the CFT, a charged scalar operator condenses.

A simple bulk action has one great virtue, namely a kind of universality: the results may be true for many different dual CFT’s, independent of the details of their dynamics. For the Einstein-Maxwell-scalar case, a fruitful exercise is to study various functional forms for the scalar potential and to scan through values of couplings in that potential Franco:2009yz (); Franco:2009if (); Aprile:2009ai (). Generally speaking, scanning through values of these parameters corresponds to scanning through many different dual CFT’s. As shown in refs. Franco:2009yz (); Franco:2009if (); Aprile:2009ai (), such changes can have a dramatic effect, for example, the phase transition can change from second to first order.

AdS/CFT can also describe superfluid states in which the condensing operator is a vector and hence rotational symmetry is broken, that is, p-wave superfluid states Gubser:2008wv (); Herzog:2009ci (). Here the CFT has a global symmetry and hence three conserved currents , where label the generators of . For a sufficiently large charge density for some subgroup of , say a sufficiently large , holographic calculations reveal that, of the known available states, those with lowest free energy have a nonzero . Not only is the broken, but spatial rotational symmetry is also broken to some subgroup.

On the AdS side, a simple bulk action that can describe such a transition is Einstein-Yang-Mills theory with gauge group . CFT states with nonzero are dual to black hole solutions with nonzero vector field in the bulk, where is the radial coordinate of AdS space. States with nonzero are dual to black hole solutions with a nontrivial . The superfluid phase transition is thus dual to charged AdS black holes developing vector hair. A string theory realization for this model is given in refs. Ammon:2008fc (); Basu:2008bh (); Ammon:2009fe (); Peeters:2009sr ().

Unlike the Einstein-Maxwell-scalar case, Einstein-Yang-Mills theory has only a single free parameter, , where is the gravitational constant (we will work in dimensions, hence the subscript) and is the Yang-Mills coupling. The Yang-Mills source terms on the right-hand-side of Einstein’s equation are proportional to . To date, most analyses of the holographic p-wave superfluid transition have employed the so-called probe limit, which consists of taking so that the gauge fields have no effect on the geometry, which becomes simply AdS-Schwarzschild. The probe limit was sufficient to show that a superfluid phase transition occurs and is second order.

Our goal is to study the effect of finite , that is, to study the back-reaction of the gauge fields on the metric. We will work with -dimensional Einstein-Yang-Mills theory, with finite . We will numerically construct asymptotically AdS charged black hole solutions with vector hair (for similar studies see refs. Gubser:2008zu (); Basu:2009vv (); Manvelyan:2008sv ()). Our principal result is that for a sufficiently large value of the phase transition becomes first order. More specifically, we find a critical value in units of the AdS radius, such that the transition is second order when and first order when .

We can provide some intuition for what increasing means, in CFT terms, as follows. Generically, in AdS/CFT , where is the central charge of the CFT Henningson:1998gx (); Balasubramanian:1999re (), which, roughly speaking, counts the total number of degrees of freedom in the CFT. Correlation functions involving the current will generically be proportional to Herzog:2009xv (); Hartnoll:2009sz (); McGreevy:2009xe (). We may (again roughly) think of as counting the number of degrees of freedom in the CFT that carry charge. Intuitively, then, in the CFT increasing means increasing the ratio of charged degrees of freedom to total degrees of freedom.

The paper is organized as follows. In section 2 we write the action of our model and discuss our ansatz for the bulk fields. In section 3 we describe how to extract thermodynamic information from our solutions. In section 4 we present numerical results demonstrating that increasing changes the order of the phase transition. We conclude with some discussion and suggestions for future research in section 5.

2 Holographic Setup

We consider Einstein-Yang-Mills theory in -dimensional asymptotically AdS space. The action is

(1)

where is the five-dimensional gravitational constant, is the cosmological constant, with being the AdS radius, and is the Yang-Mills coupling constant. The field strength is

(2)

where , with being the AdS radial coordinate, and is the totally antisymmetric tensor with . The are the components of the matrix-valued gauge field, , where the are the generators, which are related to the Pauli matrices by . includes boundary terms that do not affect the equations of motion, namely the Gibbons-Hawking boundary term as well as counterterms required for the on-shell action to be finite. We will write explicitly in section 3.

The Einstein and Yang-Mills equations derived from the above action are

(3)
(4)

where the Yang-Mills stress-energy tensor is

(5)

Following ref. Gubser:2008wv (), to construct charged black hole solutions with vector hair we choose a gauge field ansatz

(6)

The motivation for this ansatz is as follows. In the field theory we will introduce a chemical potential for the symmetry generated by . We will denote this as . The bulk operator dual to the density is , hence we include in our ansatz. We want to allow for states with a nonzero , so in addition we introduce . With this ansatz for the gauge field, the Yang-Mills stress-energy tensor in eq. (5) is diagonal. Solutions with nonzero will preserve only an subgroup of the rotational symmetry, so our metric ansatz will respect only . Furthermore, given that the Yang-Mills stress-energy tensor is diagonal, a diagonal metric is consistent. We will also pattern our metric ansatz after the ones used in ref. Manvelyan:2008sv () since these tame singular points in the equations of motion. Our metric ansatz is

(7)

with . For our black hole solutions we will denote the position of the horizon as . The AdS boundary will be at .

Inserting our ansatz into the Einstein and Yang-Mills equations yields five equations of motion for and one constraint equation from the component of the Einstein equations. The dynamical equations can be recast as (prime denotes )

(8)

The equations of motion are invariant under four scaling transformations (invariant quantities are not shown),

where in each case is some real positive number. Using (I) and (II) we can set the boundary values of both and to one, so that the metric will be asymptotically AdS. We are free to use (III) to set to be one, but we will retain as a bookkeeping device. We will use (IV) to set the AdS radius to one.

A known analytic solution of the equations of motion is an asymptotically AdS Reissner-Nordström black hole, which has , , , and , where and . Here is the value of at the boundary, which in CFT terms is the chemical potential.

To find solutions with nonzero we resort to numerics. We will solve the equations of motion using a shooting method. We will vary the values of functions at the horizon until we find solutions with suitable values at the AdS boundary. We thus need the asymptotic form of solutions both near the horizon and near the boundary .

Near the horizon, we define and then expand every function in powers of with some constant coefficients. Two of these we can fix as follows. We determine by the condition , which gives that . Additionally, we must impose for to be well-defined as a one-form (see for example ref. Kobayashi:2006sb ()). The equations of motion then impose relations among all the coefficients. A straightforward exercise shows that only four coefficients are independent,

(9)

where the subscript denotes the order of (so is the value of at the horizon, etc.). All other near-horizon coefficients are determined in terms of these four.

Near the boundary we define and then expand every function in powers of with some constant coefficients. The equations of motion again impose relations among the coefficients. The independent coefficients are

(10)

where here the subscript denotes the power of . All other near-boundary coefficients are determined in terms of these.

We used scaling symmetries to set . Our solutions will also have since we do not want to source the operator in the CFT ( will be spontaneously broken). In our shooting method we choose a value of and then vary the four independent near-horizon coefficients until we find a solution which produces the desired value of and has and .

In what follows we will often work with dimensionless coefficients by scaling out factors of . We thus define the dimensionless functions , and , while and are already dimensionless.

3 Thermodynamics

Next we will describe how to extract thermodynamic information from our solutions. Our solutions describe thermal equilibrium states in the dual CFT. We will work in the grand canonical ensemble, with fixed chemical potential .

We can obtain the temperature and entropy from horizon data. The temperature is given by the Hawking temperature of the black hole,

(11)

Here is the surface gravity of the black hole, with being the norm of the timelike Killing vector, and in the second equality we write in terms of near-horizon coefficients. In what follows we will often convert from to simply by inverting the above equation. The entropy is given by the Bekenstein-Hawking entropy of the black hole,

(12)

where denotes the area of the horizon and .

The central quantity in the grand canonical ensemble is the grand potential . In AdS/CFT we identify with times the on-shell bulk action in Euclidean signature. We thus analytically continue to Euclidean signature and compactify the time direction with period . We denote the Euclidean bulk action as and as its on-shell value (and similarly for other on-shell quantities). Our solutions will always be static, hence will always include an integration over the time direction, producing a factor of . To simplify expressions, we will define . Starting now, we will refer to as the action. includes a bulk term, a Gibbons-Hawking boundary term, and counterterms,

(13)

and exhibit divergences, which are canceled by the counterterms in . To regulate these divergences we introduce a hypersurface with some large but finite . We will always ultimately remove the regulator by taking . Using the equations of motion, for our ansatz is

(14)

For our ansatz, the Euclidean Gibbons-Hawking term is

(15)

where is the induced metric on the hypersurface and is the outward-pointing normal vector. The only divergence in comes from the infinite volume of the asymptotically AdS space, hence, for our ansatz, the only nontrivial counterterm is

(16)

Finally, is related to the on-shell action, , as

(17)

The chemical potential is simply the boundary value of . The charge density of the dual field theory can be extracted from by

(18)

Similarly, the current density is

(19)

The expectation value of the stress-energy tensor of the CFT is Balasubramanian:1999re (); deHaro:2000xn ()

(20)

where and is the extrinsic curvature. We find

(21)

Notice that . Even in phases where the current is nonzero, the fluid will have zero net momentum. Indeed, this result is guaranteed by our ansatz for the gauge field which implies a diagonal Yang-Mills stress-energy tensor and a diagonal metric (the spacetime is static).

For , , , , and we recover the correct thermodynamic properties of the Reissner-Nordström black hole, which preserves the rotational symmetry. For example, we find that and . For solutions with nonzero , the is broken to . In these cases, we find that . Just using the equations above, we also find . In the superfluid phase, both the nonzero and the stress-energy tensor indicate breaking of .

Tracelessness of the stress-energy tensor (in Lorentzian signature) implies , which is indeed true for eq. (21), so in the dual CFT we always have a conformal fluid. The only physical parameter in the CFT is thus the ratio .

4 Phase Transitions

In this section we present our numerical results. We scanned through values of from to . Typical solutions for the metric and gauge field functions appear in Figure 1. The solutions for other values of are qualitatively similar. Notice that all boundary conditions are met: at the horizon vanishes, and at the boundary and .

Figure 1: (a) The dimensionless gauge field components (red dashed) and (green dot-dashed) and the dimensionless metric function , scaled down by a factor of , (black solid) versus the AdS radial coordinate for at . (b) The dimensionless metric functions (red dashed) and (green dot-dashed) versus the AdS radial coordinate for at .

For every value of that we use, we find Reissner-Nordström solutions for all temperatures, and for sufficiently low temperatures we always find additional solutions, with nonzero , that are thermodynamically preferred to the Reissner-Nordström solution. In other words, for every value of that we use, we find a phase transition, at some temperature , in which a charged black hole grows vector hair, which in the CFT is a p-wave superfluid phase transition. Our numerical results show that the phase transition is second order for and first order for where .

For example, for , we only find solutions with until a temperature where a second set of solutions, with nonzero , appears. Figure 2 shows that rises continuously from zero as we decrease below .

Figure 2: The order parameter , multiplied by , versus the rescaled temperature for different : (green dotted), (blue solid) and (red dashed). The black dot-dashed curve is the function with . The green dotted curve is scaled up by a factor of while the red dashed curve is scaled down by a factor of such that , which depends on , coincides for the green dotted and blue solid curves. If we decrease toward , entering the figure from the right, we see that the blue solid and the green dotted curves rise continuously and monotonically from zero at , signaling a second-order phase transition. The close agreement with the black dot-dashed curve suggests that these grow from zero as . In the case, the red dashed curve becomes multi-valued at . In this case, at , the value of jumps from zero to the upper part of the red dashed curve, signaling a first-order transition.

Figure 3 (a) shows the grand potential , divided by , versus the rescaled temperature for . The blue solid curve in Figure 3 (a) comes from solutions with and the red dashed curve comes from solutions with .

Figure 3: (a) versus the rescaled temperature for . The blue solid curve comes from solutions with while the red dashed curve comes from solutions with nonzero . For , we have only the blue curve, but when the red dashed curve appears and has the lower , indicating a phase transition at . is continuous and differentiable at . (b) versus for . The blue solid and red dashed curves have the same meaning as in (a). is continuous but not differentiable at , indicating a second-order transition.

We see clearly that at the states with have the lower and hence are thermodynamically preferred. We thus conclude that a phase transition occurs at . The nonzero indicates spontaneous breaking of and of rotational symmetry down to , and hence is an order parameter for the transition. Figure 3 (b) shows the entropy , divided by , versus the rescaled temperature for . The blue solid curve and the red dashed curve have the same meaning as in Figure 3 (a). Here we see that is continuous but has a kink, i.e. a discontinuous first derivative, clearly indicating a second-order transition. For other values of , the figures are qualitatively similar.

A good question concerning these second-order transitions is: what are the critical exponents? In the probe limit, , an analytic solution for the gauge fields exists for near Basu:2008bh (), which was used in ref. Herzog:2009ci () to show that for , . In other words, in the probe limit the critical exponent for takes the mean-field value . Does increasing change the critical exponent? Our numerical evidence suggests that the answer is no: for all , we appear to find (see Figure 2).

As increases past , we see a qualitative change in the thermodynamics. Consider for example . Here again we only find solutions with down to some temperature where two new sets of solutions appear, both with nonzero . In other words, three states are available to the system: one with and two with nonzero . Figure 2 shows that as we cool the system, becomes multi-valued at . To determine which state is thermodynamically preferred, we compute the grand potential . Figure 4 (a) shows versus . The blue solid curve and the red dashed curve have the same meanings as in Figure 3.

Figure 4: (a) versus the rescaled temperature for . The blue solid and red dashed curves have the same meanings as in Figure 3. For we have only the blue solid curve. At , the red dashed curve appears and becomes multi-valued. When the red dashed curve has the lowest , indicating a phase transition at . is continuous but not differentiable at , signaling a first-order transition. (b) versus for . is not continuous at , but rather jumps from the blue solid curve to the lowest branch of the red dashed curve, indicating a first-order transition.

We immediately see the characteristic “swallowtail” shape of a first-order phase transition. If we decrease , entering the figure along the blue solid curve from the right, we reach the temperature where the new solutions appear (as the red dashed curve). The blue solid curve still has the lowest until (by definition). If we continue reducing below , then the red curve has the lowest . The transition is clearly first order: has a kink at . We can also see from the entropy that the transition is first order. Figure 4 (b) shows versus . The entropy, like the grand potential, is multi-valued, and jumps discontinuously from the blue solid curve to the lowest part of the red dashed curve at , indicating a first-order transition.

Notice that a crucial difference between (second order) and (first order) is that for the critical temperature is not simply the temperature at which becomes nonzero. We need more information to determine when , for example we can study .

A good question is: how does increasing change ? Table 1 shows several values of and the associated in units of fixed . In the probe limit, , we have the analytic result from ref. Herzog:2009ci () that . For finite , the general trend is that decreases as we increase .

Table 1: Values of , scaled up by a factor of , for various values of . Recall that the critical value of is .

5 Discussion and Outlook

We studied asymptotically AdS charged black holes in -dimensional Einstein-Yang-Mills theory with finite , that is, with back-reaction of the gauge fields. Our numerical solutions show that, for a given value of , as the temperature decreases the black holes grow vector hair. Via AdS/CFT, this process appears as a phase transition to a p-wave superfluid state in a strongly-coupled CFT. We have shown that the order of the phase transition depends on the value of : for values below , the transition is second order, while for larger values the transition is first order.

As we mentioned in the introduction, intuitively we may think of increasing as increasing the ratio of charged degrees of freedom to total degrees of freedom in the CFT. To make that intuition precise, we can consider a specific system. One string theory realization of gauge fields in AdS space is type IIB supergravity in -dimensional AdS space (times a five-sphere) plus two coincident D7-branes that provide the gauge fields Ammon:2008fc (); Basu:2008bh (); Ammon:2009fe (); Peeters:2009sr (). The dual field theory is supersymmetric Yang-Mills theory, in the limits of large and large ’t Hooft coupling, coupled to a number of massless supersymmetric hypermultiplets in the representation of , i.e. flavor fields. The global is an isospin symmetry. Translating from gravity to field theory quantities, we have and , hence , which supports our intuition. We must be cautious, however. In the field theory, the probe limit consists of neglecting quantum effects due to the flavor fields because these are suppressed by powers of . If becomes finite, then, for example, in the field theory the coupling would run, the dual statement being that in type IIB supergravity the dilaton would run, which is an effect absent in our model. We should not draw too close an analogy between our simple model and this particular string theory system.

Returning to our simple system, with fully back-reacted solutions we can potentially answer many questions:

What happens as we increase further? As increases, appears to decrease. An obvious question is whether ever becomes zero. The second-order transition occurs because of an instability of the gauge field in the Reissner-Nordström background Gubser:2008wv (). At , that instability only exists for values of below some upper bound Basu:2009vv (). If the transition was always second-order, then we would conclude that the transition is only possible for below the bound: if is greater than the bound, then the instability never appears, even if we cool the system to . If the transition becomes first-order, however, then we must rethink the bound: now Reissner-Nordström becomes metastable, and a phase transition occurs, at temperatures above those where the instability appears, so now may go to zero for some above the bound (if at all).

Does a new scaling symmetry emerge at zero temperature? The zero-temperature analysis of ref. Basu:2009vv () suggests that the metric exhibits a new, emergent, scaling symmetry. Our preliminary numerical results suggest that indeed, in the approach to zero temperature, an emergent scaling symmetry appears, of the form suggested in ref. Basu:2009vv (). The limit is numerically challenging, however, so a firm answer must wait.

What about the transport properties of the dual conformal fluid, for example the electrical conductivity, which at zero temperature should exhibit a “hard gap,” as explained in ref. Basu:2009vv ()? What is the fluid’s response to nonzero superfluid velocities? In similar systems, sufficiently large superfluid velocities also changed the transition from second to first order Basu:2008st (); Herzog:2008he (). What is the speed of sound, which need not be the same in all directions since rotational symmetry is broken, or the speeds of second and fourth sounds Herzog:2009ci (); Yarom:2009uq (); Herzog:2009md ()?

We plan to investigate these and related questions in the future.

Acknowledgements

We are grateful to A. Buchel and M. Rangamani for discussions. This work was supported in part by The Cluster of Excellence for Fundamental Physics - Origin and Structure of the Universe. M. A. would also like to thank the Studienstiftung des deutschen Volkes for financial support.

References

  • (1) J. M. Maldacena, The large N limit of superconformal field theories and supergravity, Adv. Theor. Math. Phys. 2 (1998) 231–252, arXiv:hep-th/9711200.
  • (2) C. P. Herzog, Lectures on Holographic Superfluidity and Superconductivity, J. Phys. A42 (2009) 343001, arXiv:0904.1975.
  • (3) S. A. Hartnoll, Lectures on holographic methods for condensed matter physics, Class. Quant. Grav. 26 (2009) 224002, arXiv:0903.3246.
  • (4) J. McGreevy, Holographic duality with a view toward many-body physics, arXiv:0909.0518.
  • (5) E. Witten, Anti-de Sitter space, thermal phase transition, and confinement in gauge theories, Adv. Theor. Math. Phys. 2 (1998) 505–532, arXiv:hep-th/9803131.
  • (6) J. L. Davis, P. Kraus, and A. Shah, Gravity Dual of a Quantum Hall Plateau Transition, JHEP 11 (2008) 020, arXiv:0809.1876.
  • (7) D. T. Son, Toward an AdS/cold atoms correspondence: a geometric realization of the Schroedinger symmetry, Phys. Rev. D78 (2008) 046003, arXiv:0804.3972.
  • (8) K. Balasubramanian and J. McGreevy, Gravity duals for non-relativistic CFTs, Phys. Rev. Lett. 101 (2008) 061601, arXiv:0804.4053.
  • (9) S.-J. Rey, String theory on thin semiconductors: Holographic realization of Fermi points and surfaces, Prog. Theor. Phys. Suppl. 177 (2009) 128–142, arXiv:0911.5295.
  • (10) H. Liu, J. McGreevy, and D. Vegh, Non-Fermi liquids from holography, arXiv:0903.2477.
  • (11) M. Cubrovic, J. Zaanen, and K. Schalm, Fermions and the AdS/CFT correspondence: quantum phase transitions and the emergent Fermi-liquid, arXiv:0904.1993.
  • (12) S. S. Gubser, Breaking an Abelian gauge symmetry near a black hole horizon, Phys. Rev. D78 (2008) 065034, arXiv:0801.2977.
  • (13) S. A. Hartnoll, C. P. Herzog, and G. T. Horowitz, Building a Holographic Superconductor, Phys. Rev. Lett. 101 (2008) 031601, arXiv:0803.3295.
  • (14) S. A. Hartnoll, C. P. Herzog, and G. T. Horowitz, Holographic Superconductors, JHEP 12 (2008) 015, arXiv:0810.1563.
  • (15) S. Franco, A. Garcia-Garcia, and D. Rodriguez-Gomez, A general class of holographic superconductors, arXiv:0906.1214.
  • (16) S. Franco, A. M. Garcia-Garcia, and D. Rodriguez-Gomez, A holographic approach to phase transitions, arXiv:0911.1354.
  • (17) F. Aprile and J. G. Russo, Models of Holographic superconductivity, arXiv:0912.0480.
  • (18) S. S. Gubser and S. S. Pufu, The gravity dual of a p-wave superconductor, JHEP 11 (2008) 033, arXiv:0805.2960.
  • (19) C. P. Herzog and S. S. Pufu, The Second Sound of SU(2), JHEP 04 (2009) 126, arXiv:0902.0409.
  • (20) M. Ammon, J. Erdmenger, M. Kaminski, and P. Kerner, Superconductivity from gauge/gravity duality with flavor, Phys. Lett. B680 (2009) 516–520, arXiv:0810.2316.
  • (21) P. Basu, J. He, A. Mukherjee, and H.-H. Shieh, Superconductivity from D3/D7: Holographic Pion Superfluid, JHEP 11 (2009) 070, arXiv:0810.3970.
  • (22) M. Ammon, J. Erdmenger, M. Kaminski, and P. Kerner, Flavor Superconductivity from Gauge/Gravity Duality, JHEP 10 (2009) 067, arXiv:0903.1864.
  • (23) K. Peeters, J. Powell, and M. Zamaklar, Exploring colourful holographic superconductors, JHEP 09 (2009) 101, arXiv:0907.1508.
  • (24) S. S. Gubser, Colorful horizons with charge in anti-de Sitter space, Phys. Rev. Lett. 101 (2008) 191601, arXiv:0803.3483.
  • (25) P. Basu, J. He, A. Mukherjee, and H.-H. Shieh, Hard-gapped Holographic Superconductors, arXiv:0911.4999.
  • (26) R. Manvelyan, E. Radu, and D. H. Tchrakian, New AdS non Abelian black holes with superconducting horizons, Phys. Lett. B677 (2009) 79–87, arXiv:0812.3531.
  • (27) M. Henningson and K. Skenderis, The holographic Weyl anomaly, JHEP 07 (1998) 023, arXiv:hep-th/9806087.
  • (28) V. Balasubramanian and P. Kraus, A stress tensor for anti-de Sitter gravity, Commun. Math. Phys. 208 (1999) 413–428, arXiv:hep-th/9902121.
  • (29) S. Kobayashi, D. Mateos, S. Matsuura, R. C. Myers, and R. M. Thomson, Holographic phase transitions at finite baryon density, JHEP 02 (2007) 016, arXiv:hep-th/0611099.
  • (30) S. de Haro, S. N. Solodukhin, and K. Skenderis, Holographic reconstruction of spacetime and renormalization in the AdS/CFT correspondence, Commun. Math. Phys. 217 (2001) 595–622, arXiv:hep-th/0002230.
  • (31) P. Basu, A. Mukherjee, and H.-H. Shieh, Supercurrent: Vector Hair for an AdS Black Hole, Phys. Rev. D79 (2009) 045010, arXiv:0809.4494.
  • (32) C. P. Herzog, P. K. Kovtun, and D. T. Son, Holographic model of superfluidity, Phys. Rev. D79 (2009) 066002, arXiv:0809.4870.
  • (33) A. Yarom, Fourth sound of holographic superfluids, JHEP 07 (2009) 070, arXiv:0903.1353.
  • (34) C. P. Herzog and A. Yarom, Sound modes in holographic superfluids, arXiv:0906.4810.
Comments 0
Request Comment
You are adding the first comment!
How to quickly get a good reply:
  • Give credit where it’s due by listing out the positive aspects of a paper before getting into which changes should be made.
  • Be specific in your critique, and provide supporting evidence with appropriate references to substantiate general statements.
  • Your comment should inspire ideas to flow and help the author improves the paper.

The better we are at sharing our knowledge with each other, the faster we move forward.
""
The feedback must be of minimum 40 characters and the title a minimum of 5 characters
   
Add comment
Cancel
Loading ...
247984
This is a comment super asjknd jkasnjk adsnkj
Upvote
Downvote
""
The feedback must be of minumum 40 characters
The feedback must be of minumum 40 characters
Submit
Cancel

You are asking your first question!
How to quickly get a good answer:
  • Keep your question short and to the point
  • Check for grammar or spelling errors.
  • Phrase it like a question
Test
Test description