#### Terminal holographic complexity

HJE
Terminal holographic complexity
J.L.F. Barbo´n 0 1
J. Mart´ın-Garc´ıa 0 1
0 C/ Nicol ́as Cabrera 13 , 28049 Madrid , Spain
1 Instituto de F ́ısica Teo ́rica, Universidad Auto ́noma de Madrid, IFT-UAM/CSIC
We introduce a quasilocal version of holographic complexity adapted to 'terminal states' such as spacelike singularities. We use a modification of the action-complexity ansatz, restricted to the past domain of dependence of the terminal set, and study a number of examples whose symmetry permits explicit evaluation, to conclude that this quantity enjoys monotonicity properties after the addition of appropriate counterterms. A notion of 'complexity density' can be defined for singularities by a coarse-graining procedure. This definition assigns finite complexity density to black hole singularities but vanishing complexity density to either generic FRW singularities or chaotic BKL singularities. We comment on the similarities and differences with Penrose's Weyl curvature criterion.
Spacetime Singularities; AdS-CFT Correspondence; Black Holes; Gauge-
1 Introduction
2
3
4
5
1
The local component of the terminal complexity
Local terminal complexity and coarse-graining
Terminal monotonicity
Conclusions and outlook
Introduction
analysis of certain cosmological singularities with controlled AdS/CFT embedding. Here
we seek to provide quasilocal notions of complexity which may be abstracted from
particular AdS/CFT constructions, and therefore having a larger degree of generality. While we
use the volume-complexity (VC) proposal [
3
] as a heuristic guide, most of our discussion
is tailored to the more covariant action-complexity (AC) proposal [
4, 5
].
– 1 –
The connection between spacetime singularities and complexity has a long history,
going back to the occurrence of classical chaos in generic cosmological singularities [
6–9
]
(see [10] for a recent review). In [11] Penrose gave a local criterion for the complexity
of a singularity. The basic observation is that ‘ordered’ singularities, such as those
arising in FRW models, have vanishing Weyl curvature, whereas more generic ones, such as
those arising in gravitational collapse, have a generically divergent Weyl tensor. Penrose
argued that the Weyl criterion would be associated to a large gravitational entropy flowing
into the singularity, a suggestion based on the heuristic picture of a generic cosmological
crunch, full of chaotic black hole collisions. Since black holes are known to carry entropy, a
corresponding notion of entropy may be assigned to the union of all singularities enclosed
by the colliding black holes.
One basic observation of this paper is that a suitable version of holographic complexity,
rather than entropy, provides a more natural measure of complexity of a singularity. The
proposal uses a restriction of the standard AC ansatz of [
4, 5
] to the causal domain of
dependence of the singularity. Furthermore, we will see that a local notion of complexity,
different from Penrose’s Weyl curvature, can be naturally introduced in the holographic
formalism through a coarse-graining procedure.
The paper is organized as follows. In section 2 we lay down some general definitions of
‘terminal holographic complexity’. In section 3 we discuss the local contribution to
complexity and the coarse-graining procedure. In section 4 we investigate the monotonicity
properties of this quantity in some examples of singularities admitting a completely
analytic treatment. Finally, in section 5 we offer some conclusions and outlook for plausible
generalizations.
2
A quasilocal AC ansatz for terminals
A rather intuitive notion of holographic complexity is provided by the VC ansatz of [
3
].
Formally, it looks like a generalization of the HRT construction for holographic entanglement
entropy [12, 13], removing one unit of codimension. One considers extremal
codimensionone surfaces anchored on boundary holographic data, and their volume, in an appropriate
normalization, yields the VC complexity of the dual state (see also [14]). A more
covariant prescription (AC complexity) was subsequently introduced in [
4, 5
], where one is
instructed to integrate the bulk classical action over the full causal domain of dependence
of the extremal surfaces, henceforth referred as the Wheeler-de Witt (WdW) patch.
The basic physical guide is that either ansatz provides a linear growth of quantum
complexity for a high-temperature CFT thermofield double or, in the dual picture, a large
eternal AdS black hole,
where S, T, M denote entropy, temperature and mass, respectively. This law is supposed to
apply for t ≫ T −1, up to O(1) coefficients. The detailed dependence on these coefficients
is argued to be more uniform for the AC prescription, although the physics is qualitatively
dC
dt
∼ T S ∼ M ,
– 2 –
(2.1)
complexity ‘flowing’ into the black-hole singularity S . On the right, the subtracted
codimensionone surface St′ which accounts for the time-dependence of VC complexity in the eternal black hole
∗
geometry.
HJEP06(218)3
the same, at least for large enough temperatures.1 Still, some qualitative differences in the
AC/VC dichotomy persist, particularly for cold systems, such as near-extremal black holes
or cold hyperbolic black holes. This is testimony of our still quite poor understanding of
the duality [
18–20
].
In the benchmark model provided by the eternal black hole spacetime, the central
object of interest for the VC ansatz is the extremal codimension-one surface S∞ shown in
figure 1. This surface maximizes the volume locally and it lies entirely within the black
hole interior, i.e. the past causal domain of the singularity.
The growth of complexity within the VC ansatz can be seen as the result of gradually
accessing an increasing portion of S∞. More precisely, the portion of the extremal surface
St which has a significant contribution to time dependence can be analyzed approximately
as composed of two parts: a subset of S∞ with volume proportional to t, and a transition
surface at the horizon, whose contribution to the complexity is of order S, the entropy of
the black hole. Let us denote by St′ this, loosely defined, ‘subtracted’ surface as indicated
in figure 1.
Once we decide to focus on St′ and its asymptotic limit S∞, we may consider versions
of these quantities for any terminal set S
∗ (which may in particular be a proper subset
of a wider one). The reason is that the analog of S∞ always exists given any spacelike
terminal set S∗ and its associated past domain of dependence D−(S∗) (see figure 2). Since
the volume is positive and the past boundary of D−(S∗) is null, the extremal surface is
either a local maximum of volume or it coincides with S∗ in a degenerate case. The first
situation occurs when S∗ is a standard singularity of the kind we encounter at black holes
and cosmological crunches in General Relativity, since the volume of spatial slices vanishes
at such singularities.
A more covariant version of S∞ and St′ could be obtained by adapting the AC ansatz
to this situation (cf. figure 3). Since S∞ is the extremal surface on D−(S∗), the natural
AC version of the full terminal complexity of the set S∗ is the on-shell action
C[S∗] ∝ I D−(S∗) ,
(2.2)
evaluated over the set D−(S∗). Since this definition only makes reference to the terminal
set S
∗ we regard this notion of complexity as ‘quasilocal’ and will often denote it as such.
1It has been conjectured that (2.1) should saturate the Lloyd bound [15]. See, however [16, 17].
– 3 –
Next, a notion of ‘time-dependence’ can be defined by considering a WdW patch
anchored roughly at the exit points of the cut-off surface St′, as indicated in figure 4. A
more invariant definition can be obtained by noticing that these WdW patches are nested
into one another as time increases. For sufficiently ‘late’ WdW patches, this ‘nesting’
extends to the intersections of the WdW patches with the singular set. This suggests that
we may use the nested singular subsets as a starting point in the definition of the WdW
nested family. To be more precise, let us pick a sequence of terminal subsets Su∗, ordered
by inclusion in the sense that
Su∗ ⊂ Su∗′ , for u < u ,
′
Wu = J −(Su∗) ∩ D−(S∗) .
and converging to the full terminal set S∗ as u → u∗, we can consider a set of WdW patches
Wu, defined as the intersection between D−(S∗) and the causal past of Su,
∗
past J −(S∗) .
For any given Wu, its Cauchy surfaces Σu have a common codimension-two boundary
Vu = ∂Σu (cf. figure 5), which would hold the ‘holographic data’ for Wu. For example,
Vu is a spatial section of the event horizon when S∗ is a black-hole singularity. Therefore,
we would like to interpret the ‘area’ of Vu in Planck units2 as a measure of the effective
number of holographic degrees of freedom ‘flowing’ into the terminal subset Su.
In defining the WdW patches Wu we may give privilege to the ‘anchors’, namely the
codimension-two sets Vu, or alternatively we may consider the nested family Su∗, as more
∗
2We henceforth refer to codimension-two volumes as ‘areas’.
– 4 –
(2.3)
(2.4)
HJEP06(218)3
The WdW patch Wt, associated to the cut-off surface St′, intersects the singularity
HJEP06(218)3
∗ (in red). The codimension-two set Vu is the (possibly disconnected) boundary of
Cauchy surfaces Σu for Wu.
fundamental. These two constructions are not completely equivalent, since the WdW
patch anchored at Vu may fail to intersect S∗ at sufficiently ‘early times’. In this paper we
are more interested in the asymptotic ‘late-time’ behavior in which Wu does have a
nontrivial boundary component at the singularity. Therefore, we tacitly adopt in what follows
the nesting construction of the WdW patches and we will often refer to the associated
complexity measures as ‘nesting complexity’.
In the AC/VC heuristic correspondence, the codimension-two surfaces Vu are the
natural analogs of the transition surface with volume of order S in figure 1. Therefore,
interpreting the u coordinate as a (null) time variable, we are led to the following definition of
nesting complexity associated to the given family of WdW patches Wu,
Cu∗ = α I[Wu] +
λ
4G
Area[Vu] ,
where α is a normalization factor, λ is an undetermined constant which sets the relative
importance of the codimension-two boundary counterterm, and I[Wu] denotes the on-shell
gravitational action, now integrated over the WdW patch Wu. The action can be written
as I = Ibulk + I∂ , separating bulk and boundary contributions. The bulk term has the
standard form
1
16πG
Z
Wu
I[Wu]bulk =
(R − 2Λ + Lm) + . . . ,
(2.5)
(2.6)
– 5 –
where the dots stand for higher-derivative corrections and Lm is the Lagrangian density
for non-gravitational degrees of freedom, out of which we have explicitly singled out the
cosmological constant.
The boundary term I∂ requires special consideration. On general grounds, it is given
by a sum of contributions from codimension-one and codimension-two components of the
boundary ∂Wu. The non-null codimension-one pieces and their joints are given by the
standard York-Gibbons-Hawking (YGH) term and a set of well understood joint
contributions (see [21] for a review). On the other hand, some formal choices are necessary in
the presence of null codimension-one pieces, and the physics behind these choices remains
somewhat unclear (see for example the considerations in [
19, 21, 22
]).
For the purposes of this paper, we make a minimal choice for I∂ in which we only retain
the YGH term for non-null codimension-one components and we drop the contributions
from codimension-one null components and their codimension-two joints. We do this while
keeping open the possibility that a further understanding of the microscopic definition of
complexity will require the specification of non-geometrical quantities in the AC rules.
Applying these rules to the WdW patches (2.4) we have an action
I[Wu] = I[Wu]bulk + I[Su∗]YGH ,
where the bulk term is given by (2.6) and the YGH term is restricted to the spacelike
singular component of the WdW patch,
Hence, the unpacked ansatz for the nesting complexity reads
Cu∗ = αI[Wu]bulk + αI[Su∗]YGH +
Area [Vu] .
may as well consider it as one more boundary contribution to the action.3 In this case we
are effectively picking out boundary components supported on S
u∗ ∪ Vu, rather than the
full ∂Wu. This instruction admits a nice topological interpretation, namely it amounts
to focusing on the intersection of the boundaries of J −(Su∗) and D−(S∗), rather than the
boundary of the intersection:
Su ∪ Vu = ∂J −(Su∗) ∩ ∂D−(S∗) ,
∗
nested sets Su.
∗
a relation that may be used to provide an invariant definition of Vu given the family of
Once this nesting complexity is defined, we can now recover the notion of ‘total
complexity flow’ into the singularity, which was loosely defined in (2.2), as the asymptotic limit
of the nesting procedure. More precisely, we have
3Incidentally, this would correspond to a very special case of the prescription introduced in [21], in which
one drops the codimiension-one null pieces and adjusts the normalization conventions of affine parameters
in an ad hoc, u-dependent manner.
C[S∗] = ul→imu∗ Cu ,
∗
– 6 –
(2.7)
(2.8)
(2.9)
(2.10)
(2.11)
The difference between Cu∗ and C[Su∗] as determined by the different domains of
analog of St′, (cf. figure 4).
It is important to notice that, when considering singular subsets Su∗, the nesting complexity
Cu∗ is different from the ‘total complexity’ C[Su∗] flowing into S∗, as shown in figure 6. In
u
other words, we regard C[S∗] as the AC-analog of S∞, (cf. figure 3), and Cu∗ as the
AC
In the rest of the paper we shall see that the definition of Cu∗ given in (2.9) satisfies two
interesting properties of a general character. The first is the existence of a local component
defined at space-like singularities which can be reached by a coarse-graining procedure
(section 3), and the second is the conjectured monotonicity of the nesting complexity Cu,
∗
to be discussed in section 4 below.
3
The local component of the terminal complexity
A remarkable property of the AC complexity prescription is the occurrence of a
quantitatively important contribution coming directly from the singularity, through the evaluation
of the YGH term. Since this is a term in the low-energy effective action, we should be
suspicious of its validity. On the other hand, we are instructed to take this contribution
seriously down to its precise dependence on coefficients, as this is crucial for the claimed
uniformity of the growth law (2.1) for AdS black holes in various dimensions, large and
small. In a similar vein, the contribution (or lack of it) of the YGH term at the singularities
is crucial for the ‘non-computing’ behavior in various systems, such as AdS black holes in
the 1/d expansion [20] and cold hyperbolic black holes [
18, 19
].
These considerations suggest that complexity is actually the piece of the holographic
dictionary which most efficiently ‘sees’ the properties of the singularities.
The YGH contribution is local and formally extensive over the singular set S .
However, the volume form is not generally defined at S∗, which makes the notion of ‘extensivity’
non-trivial. In order to elucidate this point, let us parametrize the near-terminal metric by
a Gaussian normal coordinate τ . This foliates the near-terminal spacetime into spacelike
surfaces Στ , according to the proper-time distance to S∗. In defining a metric on the Στ
slices, we extract a conventional power of the proper time according to the ansatz
∗
ds2 = −dτ 2 + (τ H)2γ/d dΣτ2 .
(3.1)
– 7 –
Here, H is an inverse-length setting a characteristic value for the expansion away from the
terminal set. In general, the d-dimensional metric dΣτ2 does not have a smooth limit as
τ → 0+, but we may choose the conventional exponent γ in such a way that its volume
form does have a smooth limit. We shall actually assume that this volume form is analytic
in τ , since this will be a property of all examples we study (it would be interesting to assess
the generality of this assumption). We will refer to such notion of volume for S∗ as the
‘comoving volume’ of the terminal set and denote its measure as dVolc.
In this notation, the YGH term in the action is computed as
I[S∗]YGH =
1
8πG τl→im0+ ∂τ (Hτ )γ Volc[Στ ] .
(3.2)
Picking the term proportional to the comoving volume Volc[S∗] of the singular set, we find
that the YGH term vanishes for γ > 1 and is infinite for γ < 1, except perhaps the case
γ = 0 where the answer depends on the possible occurrence of logarithmic terms in the
terminal expansion near τ = 0. The most interesting case is γ = 1, for which one defines
a nontrivial ‘comoving complexity density’ at the singular set, given by H/8πG.
The black hole singularity has γ = 1 and thus presents a purely local contribution
to complexity. In fact, this feature appears to be quite general. At spherically
symmetric black-hole singularities we have a vanishing Sd−1 and an expanding ‘radial’ direction.
Hence, the metric is locally of the Kasner form, i.e.
(3.3)
(3.4)
(3.5)
with a particular choice of Kasner parameters pc = 2/d for d − 1 ‘crunching’ directions
and pr = −1 + 2/d for the ‘ripping’ direction. More generally, the Kasner parameters are
restricted to satisfy the sum rules Pj pj = Pj pj2 = 1 and any such metric can be put in
the form (3.1) with γ = 1, with ‘comoving’ metric
In particular, it has a smooth comoving volume form,
as a simple consequence of the sum rule Pj pj = 1.
The γ = 1 property and the resulting non-vanishing ‘complexity density’ persist if we
let the Kasner parameters depend smoothly on the ‘longitudinal’ σj . In fact, the classic
results of refs. [
7–9
] (BKL) indicate that such a ‘generalized Kasner’ metrics furnish a good
local approximation of the near-singular region (after a slight generalization involving local
rescalings and frame rotations).
3.1
Evanescent terminal complexity
If we regard the generalized Kasner behavior as ‘generic’ we may say that spacelike
singularities tend to have a non-vanishing local complexity density. On the other hand, there
ds2 = −dτ 2 + X(Hτ )2pj dσj2 ,
d
j=1
dΣτ2 = X(τ H)2pj−2/d dσj2 .
j
d
dVolc[Στ ] = ∧j=1dσj
– 8 –
are important examples of singularities whose YGH contribution vanishes, such as those
occurring in standard FRW spacetimes.
To bring this simple point home, we can apply (3.2) to the standard FRW metric
ds2 = −dτ 2 + a(τ )2 dΣ2 ,
with a singularity at τ = 0. By construction, the non-singular comoving volume is just
given by the volume of the homogeneous and isotropic surfaces Σ, so that the complexity
exponent γ can be read off from the short-time asymptotics of the scale factor a(τ ). Since
FRW metrics require non-trivial matter degrees of freedom, we follow standard practice
and model them as a perfect fluid with squared speed of sound equal to w = p/ρ, where
p denotes the pressure and ρ the energy density. Then, we have the standard solution
ρ ad(1+w) = constant, which leads to a(τ ) ∼ τ 2/d(1+w) or, equivalently
(3.6)
(3.7)
(3.8)
(n) and
(3.9)
γFRW =
2
1 + w
.
This result implies that the only FRW singularity with a finite complexity density is the
slightly unphysical case with ‘stiff matter’, w = 1, leading to γ = 1. On the other hand, a
formally infinite contribution to the complexity density, associated to γ < 1, would require
w > 1 in the FRW context, i.e. a violation of the physical conditions on the matter degrees
of freedom.
Imposing the physical condition that the matter is strictly below the ‘stiff’ limit, w < 1,
we have γ > 1, implying a vanishing local complexity. Hence, we find that ‘ordered’
singularities of FRW type have a vanishing local contribution to holographic complexity,
just as it happened with the Weyl criterion of Penrose.
It turns out that there is an interesting twist to this story. According to the classic
BKL analysis [
7–9
], the vicinity of a generic spacelike singularity is not quite described by a
single generalized Kasner metric, but rather an oscillating regime where a series of ‘epochs’
succeed one another, each epoch being locally described by a generalized Kasner solution
of the type (3.3). The values of the Kasner parameters, pj , change from one epoch to the
next in a deterministic but chaotic manner. In addition, the frame determining the special
coordinates σj in (3.3) undergoes a rotation, and furthermore the induced volume form at
fixed τ is rescaled by a finite factor which we may absorb in the dimensionful expansion
parameter H. Hence, in the n-th epoch we have a metric
ds2|(n) = −dτ 2 + (Hnτ )2/d d2Σ(τn) ,
where d2Σ(τn) is a rotated version of (3.4) with Kasner parameters p(jn). All epochs are
described by γ = 1 metrics but, crucially, they have slightly different parameters pj
Hn. In particular, the substitution rule for the expansion parameter is
Hn+1 = (2p(rn) + 1) Hn ,
where p(rn) < 0 is the ‘ripping’ parameter of the n-th epoch. Since 2p(rn) + 1 < 1 for all n,
the series of Hn is monotonically decreasing.
– 9 –
If we compute the YGH contribution to complexity by placing a regulating surface
and taking the limit, the result of the complexity density is determined by the limit of Hn.
Namely it is proportional to
According to the analysis of [
7–9
] the truly generic singularity features an infinite number
of Kasner epochs. In this situation the product (3.10), featuring an infinite set of numbers
in the open interval (0, 1), is bound to vanish for almost all singularities. We refer to this
phenomenon as the ‘evanescence’ of the local complexity for a generic BKL singularity. We
may argue that, ultimately, a cutoff at Planck time from the singularity must be imposed
but, in any case, the complexity computed by this ansatz would have a suppression factor
determined by the number of epochs taking place until Planck time. These arguments
suggest that the generic singularity is not that different from the FRW one, and the standard
black-hole singularities are the ‘special ones’ regarding complexity.
3.2
Local terminal complexity and coarse-graining
The remarkable properties of the local YGH contribution beg the question of whether we
may be able to isolate this term in more physical terms. A natural strategy in this case
is to focus on the extensivity of the local contribution, a property not shared by the full
AC complexity. We can illustrate this point by focusing on the simpler case of vacuum
solutions.
A vacuum solution is a (d + 1)-dimensional Einstein manifold whose metric satisfies
Rµν =
2Λ
d − 1 gµν ,
with cosmological constant Λ and no matter degrees of freedom. The bulk contribution to
the on-shell action is then proportional to the spacetime volume
1
16πG
Z
X
I[X]bulk =
(R − 2Λ) =
Λ
complexity given formally by
∗ with non-vanishing complexity density, we have a full
C[S∗] = α
Λ
While the YGH term is extensive along the comoving volume of S∗, the bulk contribution is
extensive in the full spacetime volume of the past domain of dependence. Considering the
case Λ < 0, as corresponds to states in an AdS/CFT context, we have a negative-definite
bulk contribution, leading to a ‘subextensivity’ property of the full quasilocal complexity.
Indeed, under a coarse-graining of the singular set S
expression (3.13) satisfies
∗ = ∪iSi∗ as indicated in figure 7, the
C[S∗] = C [∪iSi∗] ≤ X C[Si∗] .
i
(3.11)
(3.12)
(3.13)
(3.14)
HJEP06(218)3
When the bulk action is dominated by a negative cosmological constant, the bulk
contribution to the AC terminal complexity of S∗ is subextensive. While the YGH contribution is
extensive over S∗, the bulk contribution is more negative for the larger set (yellow) than it is for
the union of the smaller sets (blue).
The inequality is reversed (corresponding to superextensivity) for vacuum singularities in
Λ > 0 spaces. The deviation from extensivity would disappear if the bulk contributions
were to become negligible, a situation we may expect in the limit of extreme coarse graining,
illustrated in figure 8.
In this case, the limit of infinite coarse-graining does isolate the YGH term. To be
∗
more precise, we require that the bulk contributions be consistently smaller than the YGH
contribution for small subsets of S .
We can check this explicitly for γ = 1 vacuum
singularities described by (3.13) and admitting a local Kasner description. Let us consider
a fine partition of the singular set by subsets Sǫ∗ with comoving volume of order ǫd. The
condition for the bulk contribution to be negligible for small sets is that
Vol [D−(Sǫ∗)]
Volc [Sǫ∗]
∼ ǫ a ,
(3.15)
condition (3.15) for B−(Sǫ∗).
with a > 0. Instead of computing the volume of the past domain of dependence, D−(Sǫ∗), it
is easier to compute the volume of the larger set B−(Sǫ∗), which ‘boxes’ it in the standard
coordinate frame. If τǫ is the maximal value of the τ coordinate in D−(Sǫ∗), the ǫ-box
is defined by the full τ ≤ τǫ subset with given comoving coordinates covering S∗, cf.
figure 9. Evidently, Vol [B−(Sǫ∗)] ≥ Vol [D−(Sǫ∗)], so that it is enough to establish the
In order to construct explicitly B−(Sǫ∗) for the Kasner metric (3.3) we define Sǫ∗ to
base S
be a d-dimensional cube in the σ coordinates with common extent Δσj = ǫ. Its comoving
volume is Volc [Sǫ∗] = ǫd and the past domain of dependence, D−(Sǫ∗), is a trapezoid with
ǫ∗ and a ridge with the topology of a (d − 1)-dimensional cube, determined by the
intersection of light rays in the spacetime plane with faster past-convergence. For any
direction σj we can define a corresponding conformal time coordinate ηj such that light
rays propagate with unit slope in the (ηj , σj ) plane. The explicit relation between ηj and
the proper time is
(1 − pj )Hηj = (Hτ )1−pj ,
(3.16)
where pj is the Kasner exponent in the direction σj . Light rays whose σj separation is ǫ
at τ = 0 converge in the past at τǫ(j) given by
Thus, the past domain of dependence of the full Sǫ∗ set is determined by the smallest τǫ(j)
or, equivalently, by the largest Kasner exponent which we denote by pc:
Hτǫ(j) =
(1 − pj )
Hτǫ =
(1 − pc)
Hǫ 1−pj
1
2
Hǫ 1−pc
1
2
When the bulk volume remains sufficiently bounded in the vicinity of the terminal
surface, the local complexity of S∗ results from the limit of an extreme coarse graining.
and squared speed of sound w = (2 − γ)/γ, the energy density scales as
ds2 = −dτ 2 + a(τ )2 dΣ2
ρ ∼ a(τ )−d(1+w) ∼ τ −2 .
With these ingredients we can compute the volume of the ǫ-box as
Vol B−(Sǫ∗) =
Z τǫ
0
dτ (Hτ ) ǫd =
(τǫH)2 ǫd ,
and verify (3.15) with a = 2(1 − pc)−1.
For solutions with matter degrees of freedom we need to check that the Lagrangian
Lm is not too singular. For instance, if we have a FRW terminal with metric
(3.17)
(3.18)
(3.19)
(3.20)
(3.21)
Approximating the action dimensionally as the volume integral of the energy density, we
estimate
I B−(Sǫ∗) bulk ∝ ǫ d Z τǫ
0
dτ τ γ 1
The physical condition that the matter equation of state remains strictly below the stiff
limit, w < 1, implies that γ > 1 and thus (3.23) is satisfied provided τǫ scales with a positive
power of ǫ. This happens for any solution which decelerates away from the singularity, since
the FRW conformal time is given by
Hη =
d
d − γ
d−γ
(Hτ ) d .
It is precisely for decelerating singularities that we have γ < d and τǫ ∼ ǫ d−γ scaling with
a positive power of ǫ, leading to an automatically well-defined coarse-graining limit.
The situation is less clear for FRW metrics that accelerate away from the singularity,
corresponding to γ ≥ d. Now the FRW conformal time plummets to −∞ as τ → 0+.
The problem in this case is that D−(Sǫ∗) is not itself well defined, as any past light cone
emanating from τ = 0 and converging at a finite value τ0 subtends an infinite comoving
volume at the terminal surface. To address this point we regularize the terminal surface
by bringing it slightly before the singularity at τ = δ, as indicated in figure (10). In other
words, we compute the past domain of dependence for a small, ǫ-sized subset of Σδ rather
than S∗. Let us denote this set Sǫδ. Its past domain of dependence, D−(Sǫδ), has an earliest
proper time which is a function of both ǫ and δ,
d
τ0(ǫ, δ) =
1
H
d−γ
(Hδ) d +
d − γ Hǫ d−γ
d
d
2
The condition for the coarse-graining procedure to be well-defined is now
(3.22)
(3.23)
(3.24)
(3.25)
(3.26)
(3.27)
as δ → 0 at fixed ǫ, since both terms vanish in the limit. Therefore, the bulk contribution
vanishes when we remove the regularization at fixed comoving volume, even before we take
ǫ → 0.
The borderline case of a fluid with a stiff matter equation of state, i.e. γ = 1, requires
a separate analysis. From (3.22) we can see that the bulk action diverges logarithmically
near the singularity, so that a regularization procedure will be needed as well. Following
the same notation as in figure 10, we get
For γ > d, this quantity vanishes linearly in δ as the terminal time cutoff is removed at
fixed ǫ. Hence, when we repeat the estimate (3.22) we find that
I hB−(Sǫδ)i
bulk
δ
∝ ǫ d Z τ0
dτ τ γ 1
τ 2 ∼ ǫd(τ0γ−1 − δ γ−1) −→ 0 ,
Volc[Sǫδ]
I B−(Sǫδ) bulk ∼ log
δ
,
whose behavior as ǫ, δ → 0 does depend on the order of the limits. For fixed regularization
parameter δ we have the expansion
which makes (3.27) approach linearly to zero as ǫ → 0. On the other hand, if we try to
remove the regularization at fixed ǫ we find
which diverges as δ → 0. Hence, the FRW singularity with stiff matter does not have a
consistent coarse-graining limit which would isolate the local complexity density.4
On the other hand, the more ‘physical’ FRW singularities with γ > 1 are extremely
‘thin’ in any measure of local complexity. For 1 < γ < d, corresponding to metrics that
decelerate away from the singularity, we have a vanishing local contribution from the YGH
term and a well-defined coarse-graining procedure isolating this term. For γ ≥ d, the
situation is even more extreme, since the quasilocal complexity vanishes for sets of small
but fixed comoving volume.
We interpret these features as holographic analogs of the Weyl curvature criterion by
Penrose, which also gave a smallest measure of complexity to FRW cosmologies. The main
difference here is that the holographic notion of complexity is rather more refined, making
a quantitative distinction between accelerating and decelerating cosmologies.
4
Terminal monotonicity
In this section we return to the full quasilocal complexity and discuss the monotonicity of its
‘nesting’ properties. Given that our definition of Cu∗ in (2.4) was tailored to mimic the role
of the cut surface St′ in the VC ansatz, we expect that Cu∗ should increase monotonically,
at least asymptotically as u → u∗. Since this may depend on the choice of ‘entropic’
coupling λ, we conjecture that an appropriate choice exists such that this always happens.
4The consistency of the coarse-graining procedure would need to be analyzed anew if extra boundary
terms are included in the action, beyond our minimal choice in (2.9) and (2.11).
In other words, we conjecture that the full on-shell action of the WdW patch I[Wu] is
either monotonically increasing as u → u∗ or, in case it decreases, it does so at a rate
bounded (in absolute magnitude) by that of Area[Vu].
A priori, it is not immediately obvious that such monotonicity will hold. Let us consider
again vacuum solutions satisfying (3.11). Given a set of WdW patches Wu associated to a
family of singular subsets S
u∗ with γ = 1, and holding codimension-two ‘edges’ Vu, we can
write the full ‘nesting’ complexity as a function of u as arising from three contributions:
∗
Cu VAC = α
Λ
The first term comes from the bulk Einstein-Hilbert action and is extensive in space-time
volume of the WdW patch. The second term is the terminal density, extensive in the
comoving volume of the terminal set, and the last term is the ‘entropic’ counterterm.
A glance at the expression (4.1) indicates that the monotonicity is not obvious, in
particular for singularities embedded in AdS, since there one finds Λ < 0 and the bulk
contribution is negative as the space-time volume of the WdW patches grows. In addition,
the particular monotonicity properties of Area[Vu] could affect the final answer. On the
other hand, the space-time volume vanishes near singularities of Einstein’s equations. Thus,
we expect the main contribution to the on-shell action to come from the ‘corners’ of the
WdW patch, i.e. the vicinity of the sets Vu. In such a situation we may expect that
any threat to monotonicity coming from the bulk action could eventually be fixed by an
appropriate choice of the entropic coupling λ.
Short of a general proof, we have examined a number of examples in which the explicit
computation can be carried reliably and found agreement with the monotonicity conjecture.
We consider here a few extreme cases which illustrate the qualitatively different roles played
by the entropic counterterm proportional to λ in (2.5). Our choices are motivated by the
ability to compute exactly the complexity on terminal WdW patches using the specific
formula (4.1), but also by our interest in exposing as much as possible the contrast between
entropy and complexity when referred to cosmological singularities. Quite generally, we
can associate an entropic measure to a terminal set S∗ by the limit of the codimension-two
areas Area[Vu] in Planck units, as u → u∗, i.e.
S[S∗] ≡ lim
u→u∗
Area[Vu] .
4G
(4.2)
With this definition, the full singularity of a standard black hole solution has finite entropy,
whereas any proper portion of the singularity has zero entropy. On the other hand, there
are prototypical cosmological singularities with infinite entropy, such as the interior of
a Coleman-de Luccia bubble in vacuum decay. It is very interesting to keep track of the
terminal complexity in these wildly different situations from the point of view of the entropy
as a measure of the ‘holographic dimensionality’ of the relevant state spaces.
Before embarking in our tour of examples, we would like to comment briefly on the
relation to previous work. In [1] the VC complexity was estimated for a number of
cosmological singularities which are naturally embedded into concrete AdS/CFT constructions.
In these examples it was found that a regularized version of the VC complexity was
monotonically decreasing on approaching the singularity, in contrast with our statement here for
the cuasilocal complexity. A similar behavior was obtained for the AC ansatz in the same
examples by [2]. The reason for this apparent discrepancy is simply that the full complexity
computed in [1] is dominated by UV contributions to the VC ansatz, and these are highly
dependent on the particular details of the embedding into asymptotically AdS geometries.
For instance, some of the examples are based on singular CFT metrics which shrink to zero
size, and others involve expanding domain walls in the bulk. In the first case it is natural
that the UV contribution to complexity should have a negative derivative in time, as
corresponds to a shrinking Hilbert space on the full CFT. In the second case, a time-dependent
conversion between UV and IR degrees of freedom is introduced in the CFT by switching
on a relevant operator with a time-dependent coupling, and the c-theorem explains why
the UV again dominates the balance. Therefore, there is no contradiction since the two
monotonicity statements refer to different quantities. The positive monotonicity of the
quasilocal complexity defined here (by restriction of the AC/VC ansatz to the interior of
D−(S∗)), is compatible with the negative monotonicity of the full complexity, particularly
when the latter is dominated by a strong UV time-dependence.
Vacuum terminals with constant entropy.
We begin our tour of examples with the
benchmark case of a (future) black hole interior. The standard case is provided by the
spherical AdS black hole solution, with an R × SO(d) isometry group and metric
dr2
f (r)
ds2 = −f (r) dt2 +
+ r2 dΩ2d−1 ,
f (r) = 1 +
r
2
ℓ2 −
µ
rd−2 ,
1
where ℓ is the curvature radius of AdS. The vicinity of the singularity at r = 0 is controlled
by a single length scale, µ d−2 , which relates to the horizon radius R through
where Ωd−1 = Vol[Sd−1]. The YGH contribution to the on-shell action is
I[SΔ∗t]YGH =
Volc[SΔ∗t] =
H
8πG
d µ Ωd−1 Δt
16πG
µ = Rd−2 +
H =
2
d · µ
Rd
ℓ2 .
1
d−2
.
The near-terminal metric can be written in the form (3.1) with γ = 1 and
The terminal set at r = 0 has topology R × Sd−1, and its comoving metric degenerates
through a stretching of the R factor and a contraction of the sphere. We can parametrize
the terminal set by the homogeneous t-coordinate along the R factor. Let SΔt denote a
subset IΔt × Sd−1 of the terminal set, where IΔt ⊂ R is an interval of length Δt along the
∗
t coordinate. Its comoving volume is then given by
Volc[SΔ∗t] = Δt Ωd−1 H 1−d ,
(4.3)
(4.4)
(4.5)
(4.6)
(4.7)
for both small and large black holes. Expressed in terms of physical quantities, we find
a contribution to complexity which is positive and proportional to Δt M , where M is the
mass of the black hole.
The bulk contribution (3.12) is negative-definite. In order to compute it, we follow [21]
and introduce an infalling Eddington-Finkelstein coordinate
u = t +
Z r dr′
f (r′) ,
to write the metric in the form
This allows us to compute the integral (3.12) as
ds2 = −f (r) du2 + 2dudr + r2dΩ2d−1 .
I[WΔt]bulk = −
Ωd−1d Z
8πGℓ2
WΔt
du dr rd−1 .
To show monotonicity, it suffices to evaluate the action over a stripe of u-extent δu, as
indicated in figure 11, and we obtain
I[δW]bulk = −
Ωd−1d Z
8πGℓ2 δu
du
Z R
0
dr rd−1 = −
Ωd−1 Rd
8πGℓ2 δu .
Adding the contribution from the YGH term and using the relation µ = Rd−2 + Rd/ℓ2, we
find that
δI = δu Ωd−1 Rd−2
8πGℓ2
d ℓ2 + (d − 1)R2
is always positive, implying the positive monotonicity of CΔ∗t. In establishing this result,
the ‘entropic’ counterterm proportional to λ plays no role, since the horizon area is
asymptotically constant.
(4.8)
(4.9)
(4.10)
(4.11)
(4.12)
Vacuum terminals with divergent entropy.
A qualitatively different situation is
obtained for terminals that look like standard cosmologies of FRW type. In general, such
FRW models require non-trivial contributions from matter degrees of freedom. In order to
stay within the realm of vacuum solutions, so that we can still apply (4.1), we look at FRW
metrics arising in the interior of vacuum bubbles of Coleman-de Luccia type. A crucial
property of any such bubble is that it expands, asymptotically approaching the speed of
light, so that the area of bounding codimension-two surfaces is guaranteed to diverge, i.e.
they always have divergent entropy.
As an explicit example with an exact solution we can consider a kind of ‘topological
crunch’ spacetime (see for example [
1, 23–25
]). In this construction, we have a sort of
higher-dimensional generalization of the BTZ black hole, i.e. the metric is locally pure
singular locus. More precisely, we consider the following metric on D−(S∗),
AdSd+2, but a convenient identification by the group of integers realizes a time-dependent
compactification with topology S1 × AdSd+1, where the S1 fiber shrinks to zero size at the
ds2 = −dt2 + ℓ2 sin2(t/ℓ)dH2d + ℓ2 cos2(t/ℓ)dφ2 ,
(4.13)
where φ is an angle parametrizing the additional compact circle, dH2d is the standard unit
metric on the d-dimensional Euclidean hyperboloid and ℓ is the AdS radius of curvature.
The singularity at t = ℓπ/2, occurring when the compact circle degenerates to vanishing
size, is very mild, but enough to guarantee that the terminal set S
has zero physical volume. The comoving volume of S∗ is given by the standard volume
∗ of topology S1 × Hd
forms on S1 × Hd times ℓd+1.
This model has the nice feature of admitting a straightforward embedding into
AdS/CFT. On the exterior of D−(S∗) we can realize the SO(1, d) isometry on timelike
de Sitter hypersurfaces instead of spacelike hyperbolic ones. Hence, the model can be
embedded as a de Sitter-invariant state of a CFTd+1 defined on dSd × S1.
In order to perform the required computations, it is useful to define null coordinates
on D−(S∗) as follows. First, we introduce a radial coordinate χ over Hd and a conformal
time variable η = 2 tan−1 et/ℓ over the AdSd+1 factor, so that the metric is written in
the form
ds2 = ℓ2 sin2(t/ℓ) −dη2 + dχ2 + sinh2(χ) dΩ2d−1 + ℓ2 cos2(t/ℓ)dφ2 .
(4.14)
(4.15)
(4.16)
Next, we introduce null coordinates
and write the complete metric as
tan U = eη+χ ,
tan V = eη−χ ,
ds2 = ℓ2 sec2(U − V ) −4dU dV + sin2(U − V )dΩ2d−1 + cos2(U + V )dφ2 .
We consider a set of nested (d + 2)−dimensional WdW patches Wu bounded by the
coordinate u, where u = U − π/4 ≥ 0. The bulk action is then given by
The causal structure of the topological crunch singularity and its associated WdW
patches. The projection of Wu onto the (U, V ) plane is shown as the yellow sector in the figure,
and coincides with the domain of integration W(U, V ).
where the domain of integration W(U, V ) in the (U, V ) plane is shown in the figure 12.
Notice that Newton’s constant G has now the appropriate dimensionality for a (d +
2)dimensional spacetime, explaining the power of ℓ in the numerator. A rather explicit
expression can be obtained for the variation with respect to the u coordinate,
d
du
Ωd−1ℓd h(d − sin(2u))(tan(u + π/4))d + sin(2u)(tan(2u))di ,
and the behavior near the singularity at u∗ = π/4 is
d
du
Area[Vu] ≈ 2π Ωd−1ℓ
d
d − 1
(u∗ − u)d
A similarly explicit expression can be obtained for the variation of the YGH term:
d
du
I[Wu]bulk ≈ −
Ωd−1ℓd
2Gd
d + 2−d − 1
1
(u∗ − u)d
.
d
du
Ωd−1ℓd (tan(2u))d
where the last expression is the asymptotic limit near the singularity. We find that, for all
d ≥ 2, the sum of (4.19) and (4.20) is negative, approaching minus-infinity as u → u∗. On
the other hand, one finds that the entropic term has the same degree of divergence near
the singularity:
Area[Vu] = 2π Ωd−1ℓd (tan(u + π/4))d−1 ,
and the rate near the singuarity is given by
(4.18)
(4.19)
(4.20)
(4.21)
(4.22)
Hence, the terminal complexity of the topological crunch has the expected
monotonicity property provided we choose
λ
α
>
1
πd
This example differs from the eternal black hole in that the ‘entropy of the singularity’
lower bound in order to guarantee monotonicity of Cu.
∗
grows indefinitely, as measured by Area(Vu), and the associated coupling λ must satisfy a
An even simpler model of this kind is obtained by removing the S1 factor in (4.13).
The FRW patch of pure AdSd+1 has a coordinate singularity at t = πℓ, but we may render
it a true singularity by considering a thin-walled bubble with dSd wordvolume hitting the
boundary of AdSd+1 right at the boundary of the t = πℓ null surface (cf. figure 13). This
model represents an approximate de Sitter-invariant condensate state for a CFT on dSd
(cf. [1, 26, 27]).
Using the same null coordinates introduced in (4.15), we have a metric
HJEP06(218)3
ds2 = ℓ2 sec2(U − V ) −4dU dV + sin2(U − V )dΩ2d−1 ,
(4.24)
The singularity sits at U = π/2 in the null limit. We can now consider the set of nested
WdW patches Wu defined by 0 ≤ U ≤ u and V ≤ U . These patches do not actually touch
the singularity in the idealized null limit, as shown in figure 13, so that the monotonicity
properties of Cu∗ depend entirely on the balance between the negative-definite bulk
contribution and the volume of the codimension-two boundary sets Vu. The bulk action is
given by
dℓd−1Ωd−1 Z u
8πG
0
dU
dV
Z U
0
where the last expression is an approximation for u ≈ π/2. On the other hand, the volume
of the codimension-two set is
Area[Vu] = ℓd−1Ωd−1 tand−1(u) ≈ ℓd−1Ωd−1
1
π2 − u
d−1
.
We thus conclude that monotonicity is guaranteed provided we pick an entropy coupling
λ satisfying the inequality
λ
α
>
1
2π(d − 1)
.
In a sense, the singularity induced by a vacuum bubble in the thin-wall limit provides
the extreme case in which the monotonicity is threatened by a negative, diverging, bulk
contribution. Here we see that this negative infinity is always tamed by a sufficiently large,
but ultimately finite, entropic counterterm. It would be interesting to obtain these results
in more realistic constructions going beyond the thin-wall approximation by the inclusion
of explicit dynamical scalar fields.
d−1
,
(4.25)
(4.26)
(4.27)
with dSd worldvolume.
A vacuum terminal with vanishing entropy. In this section we consider an example
in which the entropy of the singularity, as defined by the volume of codimension-two sets
Vu, has precisely the opposite behavior to the topological crunch model, namely it vanishes
at the singular locus. Consider the Kasner metric in d + 1 dimensions
d
i=1
ds2 = −dt2 + X(tH)2pi dxi2 ,
(4.28)
the standard vacuum solution with zero cosmological constant and Rd symmetry and, as
discussed at length in the previous section, a local approximation for ‘small portions’ of
more general singularities (cf. [
7–10
]).
We recall that the coefficients pi are restricted
to satisfy Pi pi = Pi pi2 = 1, and at least one of the exponents pi must be negative,
indicating that at least one direction stretches as one approaches the singularity. The
volume of codimension-two surfaces sitting at some constant value of t always vanishes as
|t| in the t → 0 limit.
For simplicity of computations, we shall consider a family of sets S
symmetry of the metric, defined as follows: in the Rd parametrized by the coordinates xj ,
we single out one particular direction, xu, and pick an interval Iu =
this direction. The set of nested terminal sets is then defined as the ‘slabs’ S
− u2 , u of length u in
2
u∗ = Rd−1 × Iu.
The full terminal set is obtained by taking the limit u → u∗ with some finite u∗. The
causal structure of the WdW patches is shown in figure 14. The codimension-two surfaces
Vu controlling the entropic coupling sit on the surfaces t = t(u), with t(u) determined by
u∗ adapted to the
the equation
(1 − pu)(u∗ − u) = 4 t(u)1−pu ,
(4.29)
and pu is the Kasner exponent along the compact direction of the slab Su.
∗
Since (4.28) has already the form (3.1) with γ = 1, we obtain an extensive and positive
contribution from the YGH term
where we interpret the volume of Rd−1 in the sense of defining the appropriate density
along the non-compact directions. Moreover, the Ricci-flatness property of (4.28) implies
that there is no bulk contribution to the on-shell action, and we seem to obtain a monotonic
result for Cu∗ with no explicit reference to the entropic coupling λ. On the other hand, a
simple computation shows that
Area[Vu] = (1 − pu)H Vol[Rd−1] u∗ − u
,
2
which vanishes as u → u∗, giving a monotonically decreasing contribution to complexity
as soon as we have λ > 0. Therefore, in order to ensure the right monotonicity property
of Cu∗ we must assume that the entropic coupling satisfies an upper bound
λ
α
π(1 − pu)
,
independently of whether the slab’s finite interval is laid along a ripping (pu < 0) or
crunching (pu > 0) direction.
Unlike the Coleman-de Luccia bubble, the Kasner situation has a finite complexity in
the asymptotic limit u → u∗, but the entropy counterterm could make the total complexity
approach the limit from above rather than from below. If we want to ensure a positive
rate in this case we must prevent the entropic coupling from being too large, as indicated
in the bound (4.32).
5
Conclusions and outlook
In this paper we have introduced quasilocal notions of AC complexity for terminal sets such
as spacelike singularities in General Relativity. The basic idea is to build WdW patches
restricted to the past causal domain of the singular set. Holographic data are associated to
codimension-two surfaces on corresponding horizons. In principle, we can assign a notion
of entropy to a singularity by looking at the area of these codimension-two sets.
(4.30)
(4.31)
(4.32)
One of the main observations made in this note is that the on-shell action of the WdW
patches should be defined as monotonically increasing under the nesting of WdW patches,
independently of the behavior of the entropy. We have tested this idea on a number of
exactly solvable examples comprising the three qualitative behaviors with respect to the
entropy: asymptotically constant entropy like in black-hole interiors, diverging entropy
as in the interior of Coleman-de Luccia bubbles, and vanishing entropy as in the Kasner
spacetime. In all cases one can define monotonic terminal AC complexities at the price
of adjusting an additive contribution proportional to the entropy. It would be
interesting to back these checks with actual proofs, under the assumption of appropriate energy
conditions.
The YGH term evaluated at the singularity plays a special role. First, we emphasize
that it is to be taken seriously despite the fact that it is a contribution to the low-energy
effective action extrapolated beyond its natural realm of applicability. Second, we have
shown that one may isolate it as the local contribution to the complexity under an
extreme coarse-graining procedure. In this respect, the relation of the YGH term to the full
quasilocal complexity is analogous to the relation between the classical thermodynamic
entropy, obtained through coarse-graining, and the exact von Neumann entropy of a quantum
many-body state.
Another interesting property of the local complexity is its vague resemblance to
Penrose’s Weyl curvature criterion. In particular, FRW singularities are singled out by having
vanishing complexity density, according to this definition. The similarity is not perfect
though, since one can identify various differences. First, the ‘lack of complexity’ seems to
be even more severe for FRW metrics which accelerate away from the singularity.
Second, within the local description of generic singularities, as presented in the classic BKL
analysis, our ansatz assigns a vanishing complexity density to the formal infinite sequences
of chaotic Kasner ‘epochs’. Since these chaotic structures are generic in the light of the
BKL analysis, we would conclude that the local complexity density of generic spacelike
singularities is zero. In this sense, complexity would behave similarly to local gravitational
energy in General Relativity: while global and quasilocal definitions of gravitational energy
exist, any attempt at a fully local definition is doomed to failure because of the equivalence
principle. We find the parallel with complexity unveiled here quite intriguing.
A question of potential interest is the generalization of these concepts to asymptotically
de Sitter spacetimes. This is natural since the asymptotic future of de Sitter is also a ‘causal
terminal’ and one may consider the behavior of the quantities defined in this paper. In fact,
a simple check shows that every term in the quasilocal complexity ansatz becomes infinite
in this case. Both the YGH term evaluated at the asymptotic future and the bulk action
diverge. It turns out that in this case the YGH term dominates and formally gives an
infinitely negative complexity. It would be interesting to elucidate these questions further.
Acknowledgments
We would like to thank M. Alishahiha, B. Freivogel, C. Gomez, J. Maldacena, R. Myers, E.
Rabinovici, J. Martinez-Magan and K. Sousa for discussions on various aspects of
computational complexity, as well as the participants of the ‘It for Qbit Workshop’ at Bariloche
and ‘Complexity Workshop 2018’ at AEI-Postdam, where preliminary versions of this work
were presented. This work is partially supported by the Spanish Research Agency
(Agencia Estatal de Investigaci´on) through the grants IFT Centro de Excelencia Severo Ochoa
SEV-2016-0597 and FPA2015-65480-P. The work of J.M.G. is funded by Fundaci´on La
Caixa under “La Caixa-Severo Ochoa” international predoctoral grant.
Open Access.
This article is distributed under the terms of the Creative Commons
Attribution License (CC-BY 4.0), which permits any use, distribution and reproduction in
any medium, provided the original author(s) and source are credited.
[INSPIRE].
JHEP 01 (2016) 084 [arXiv:1509.09291] [INSPIRE].
[2] S. Bolognesi, E. Rabinovici and S.R. Roy, On Some Universal Features of the Holographic
Quantum Complexity of Bulk Singularities, JHEP 06 (2018) 016 [arXiv:1802.02045]
Selected Papers. “Construction of a General Cosmological Solution of the Einstein Equation
with a Time Singularity”, World Scientific, New York U.S.A. (1996), pg. 763.
[10] V. Belinski and M. Henneaux, The Cosmological Singularity, Cambridge University Press,
Cambridge U.K. (2017).
[11] R. Penrose, Singularities and time-asymmetry, in General Relativity: An Einstein Centenary
Survey, Akademie-Verlag, Berlin Germany (1979), pg. 581.
Phys. Rev. Lett. 96 (2006) 181602 [hep-th/0603001] [INSPIRE].
[12] S. Ryu and T. Takayanagi, Holographic derivation of entanglement entropy from AdS/CFT,
[13] V.E. Hubeny, M. Rangamani and T. Takayanagi, A Covariant holographic entanglement
entropy proposal, JHEP 07 (2007) 062 [arXiv:0705.0016] [INSPIRE].
[14] M. Alishahiha, Holographic Complexity, Phys. Rev. D 92 (2015) 126009 [arXiv:1509.06614]
[arXiv:1710.01175] [INSPIRE].
JHEP 01 (2017) 062 [arXiv:1610.08063] [INSPIRE].
[arXiv:1710.06415] [INSPIRE].
HJEP06(218)3
JHEP 02 (2013) 038 [arXiv:1210.7244] [INSPIRE].
[arXiv:1308.1921] [INSPIRE].
[3] L. Susskind , Entanglement is not enough, Fortsch. Phys . 64 ( 2016 ) 49 [arXiv: 1411 .0690] [4] A.R. Brown , D.A. Roberts , L. Susskind , B. Swingle and Y. Zhao , Holographic Complexity Equals Bulk Action?, Phys. Rev. Lett . 116 ( 2016 ) 191301 [arXiv: 1509 .07876] [INSPIRE].
[5] A.R. Brown , D.A. Roberts , L. Susskind , B. Swingle and Y. Zhao , Complexity, action and black holes, Phys. Rev. D 93 ( 2016 ) 086006 [arXiv: 1512 .04993] [INSPIRE].
[6] C.W. Misner , Mixmaster universe, Phys. Rev. Lett . 22 ( 1969 ) 1071 [INSPIRE].
[7] V.A. Belinsky , I.M. Khalatnikov and E.M. Lifshitz , Oscillatory approach to a singular point in the relativistic cosmology , Adv. Phys . 19 ( 1970 ) 525 .
[8] V.A. Belinsky , I.M. Khalatnikov and E.M. Lifshitz , A General Solution of the Einstein Equations with a Time Singularity, Adv . Phys. 31 ( 1982 ) 639 .
[9] V.A. Belinskii , E.M. Lifshitz and I.M. Khalatnikov , 30 Years of the Landau Institute - [15] S. Lloyd , Ultimate physical limits to computation , Nature 406 ( 2000 ) 1047 [16] W. Cottrell and M. Montero , Complexity is simple!, JHEP 02 ( 2018 ) 039 [17] D. Carmi , S. Chapman , H. Marrochio , R.C. Myers and S. Sugishita , On the Time Dependence of Holographic Complexity , JHEP 11 ( 2017 ) 188 [arXiv: 1709 .10184] [INSPIRE].
[18] J.L.F. Barbo ´n and J. Mart´ın-Garc´ıa, Holographic Complexity Of Cold Hyperbolic Black Holes , JHEP 11 ( 2015 ) 181 [arXiv: 1510 .00349] [INSPIRE].
[19] S. Chapman , H. Marrochio and R.C. Myers , Complexity of Formation in Holography, [20] J.L.F. Barbo ´n and J. Mart´ın-Garc´ıa, Holographic non-computers , JHEP 02 ( 2018 ) 181 [21] L. Lehner , R.C. Myers , E. Poisson and R.D. Sorkin , Gravitational action with null boundaries , Phys. Rev. D 94 ( 2016 ) 084046 [arXiv: 1609 .00207] [INSPIRE].
[22] A. Reynolds and S.F. Ross , Divergences in Holographic Complexity, Class. Quant. Grav. 34 ( 2017 ) 105004 [arXiv: 1612 .05439] [INSPIRE].
[23] M. Ban ˜ados, A. Gomberoff and C. Martinez , Anti-de Sitter space and black holes, Class. Quant. Grav . 15 ( 1998 ) 3575 [ hep -th/9805087] [INSPIRE].
[24] M. Ban ˜ados, Constant curvature black holes , Phys. Rev. D 57 ( 1998 ) 1068 [ gr -qc/9703040] [25] J. Maldacena and G.L. Pimentel , Entanglement entropy in de Sitter space,