Entwinement and the emergence of spacetime
Vijay Balasubramanian
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Borun D. Chowdhury
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Bartlomiej Czech
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Jan de Boer
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Open Access
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c The Authors.
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Tempe
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Arizona 85287
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U.S.A
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365 Fifth Avenue
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New York, NY 10016
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U.S.A
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Department of Physics, Stanford University
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Institute for Theoretical Physics, University of Amsterdam
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CUNY Graduate Center, Initiative for the Theoretical Sciences
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Department of Physics, Arizona State University
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David Rittenhouse Laboratories, University of Pennsylvania
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Science Park 904
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Postbus 94485, 1090 GL Amsterdam
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The Netherlands
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382 Via Pueblo Mall
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Stanford, CA 943054060
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U.S.A
It is conventional to study the entanglement between spatial regions of a quantum field theory. However, in some systems entanglement can be dominated by internal, possibly gauged, degrees of freedom that are not spatially organized, and that can give rise to gaps smaller than the inverse size of the system. In a holographic context, such small gaps are associated to the appearance of horizons and singularities in the dual spacetime. Here, we propose a concept of entwinement, which is intended to capture this fine structure of the wavefunction. Holographically, entwinement probes the entanglement shadow  the region of spacetime not probed by the minimal surfaces that compute spatial entanglement in the dual field theory. We consider the simplest example of this scenario  a 2d conformal field theory (CFT) that is dual to a conical defect in AdS3 space. Following our previous work, we show that spatial entanglement in the CFT reproduces spacetime geometry up to a finite distance from the conical defect. We then show that the interior geometry up to the defect can be reconstructed from entwinement that is sensitive to the discretely gauged, fractionated degrees of freedom of the CFT. Entwinement in the CFT is related to nonminimal geodesics in the conical defect geometry, suggesting a potential quantum information theoretic meaning for these objects in a holographic context. These results may be relevant for the reconstruction of black hole interiors from a dual field theory.

Conical defects, long geodesics and entwinement
The entanglement shadow of the conical defect geometries
Ungauging the dual description of the conical defect
Entwinement and entanglement shadows
Reconstruction of geometry from entwinement
Geometry from entanglement
Reconstructing the conical defect spacetime
Beyond conical defects
Spatial entanglement, entwinement and mutual information
Toward higher dimensions
1 Introduction 3 4 2.1
According to the RyuTakayanagi (RT) proposal [1, 2], classical geometry and quantum
entanglement are related via holographic duality. The proposal states that the entanglement
entropy of a spatial region R in the field theory is given by:
S(R) =
4G A=R
In this formula, which assumes that the bulk spacetime is static,1 the minimum is taken
over bulk surfaces, which are contained in the same spatial slice as the boundary region
R and which asymptote to the boundary of R. The RT formula relates entanglement
entropy, a nonlocal quantity in the boundary theory, to a minimal surface, which is a
nonlocal object in the bulk. Recently, we proposed a new quantity, the differential entropy,
constructed out of entanglement, that reconstructs the areas of closed surfaces in AdS
that do not asymptote to the boundary [48].2 By shrinking such closed surfaces one can
attempt to reconstruct local geometry in AdS space from purely field theoretic objects [10
14]. The relevance of boundary entanglement for such a reconstruction was first pointed
out in [9, 1517], see also [18].
1See [3] for a generalization to nonstatic spacetimes.
2This quantity was found while trying to make the proposal of [9], that areas of general surfaces in
spacetime are directly related to entanglement across them, more precise.
In order for this program to succeed, it is necessary for the union of RT minimal
surfaces to cover all of spacetime. This is barely possible in empty AdS space where the
largest RT surfaces, associated to the entanglement of half of the field theory, are necessary
to include the origin of space. But away from pure AdS, an entanglement shadow can
develop a region which is not probed by minimal surfaces and hence by conventional
spatial entanglement entropy in the dual field theory. (Our terminology is inspired by the
term causal shadow introduced by [19] to describe a region which is causally disconnected
from all the spacetime boundaries.) For example, the AdSSchwarzschild and BTZ black
holes have an entanglement shadow of thickness of order the AdS scale surrounding the
horizon [20]. This conundrum reflects a general difficulty in the AdS/CFT correspondence
of identifying field theoretic observables associated to physics in a region of size less than
one AdS volume [2128].
How can we see inside entanglement shadows? Many lines of evidence point to the idea
that recovering the local physics in such regions involves internal degrees of freedom of
the CFT that are not themselves spatially organized. Consider, for example, the role of the
matrix degrees of freedom in the SU(N ) YangMills theory dual to AdS5 [29] and the
fractionated degrees of freedom in the D1D5 string dual to AdS3 [30] in reconstructing the deep
interior of space [31], the compact dimensions of the bulk [3234], and the entropy of AdS
black holes [3539]. Such internal degrees of freedom can have energy gaps much smaller
than those dictated by the spatial size of the system, and thus represent a deep IR regime
of the field theory that will on general grounds be associated to the deep interior of AdS.3
All of this suggests that to see inside entanglement shadows we will need to consider
the entanglement of internal degrees of freedom with each other. If the Hilbert space for
these variables factorizes, we can derive a reduced density matrix for any subset of them
and compute its entanglement entropy. However, there is often an additional subtlety
typical realizations of holography involve gauge symmetries acting on the internal degrees
of freedom. In this context, how do we ask questions like How entangled is a subset of
the degrees of freedom with the rest of the theory? The challenge here is that the subset
in question may not be gauge invariant by itself. We propose to deal with this problem in
a pedestrian fashion: embed the theory in an auxiliary, larger theory, where the degrees of
freedom are not gauged, compute conventional entanglement there, then sum over gauge
copies to get a gauge invariant result. (We will deal with discrete gauge groups in this
paper where the sum over gauge copies is easy to define; for continuous gauge symmetries
consideration of an appropriate measure would be necessary.) This is not a conventional
notion of entanglement that is associated to a gauge invariant algebra of observables. Hence
we give it a new name entwinement. The justification for inventing this concept is that
it will turn out to have a useful meaning in the dual gravity theory in terms of extremal
but nonminimal surfaces in the examples we consider.4
3See [40] for a discussion of entanglement between high and low momenta in a field theory. We are
here discussing entanglement between IR degrees of freedom that are not spatially organized, so we need a
4A notion of ungauging or expanding the Hilbert space and then regauging has also appeared in [41, 42],
which discuss the problem of seeing behind a horizon in AdS/CFT.
A trivial example of entwinement is to consider two CFTs living on the same space.
One can compute the entanglement entropy of the degrees of freedom living in a spatial
region R in CFT1 and CFT2 separately, and then sum up the results. This is a rudimentary
example of entwinement. When the two CFTs are in a product state, this computation
returns the entanglement entropy of the region R, but in a general state entwinement is
distinct from entanglement. A different example is to consider matrix degrees of freedom
can imagine computing the entanglement of a subset of the matrix degrees of freedom in
some spatial region. However, because of the gauge symmetry we cannot do this navely.
One option is to ungauge and then sum over gauge copies; this would give the entwinement
that we propose to define. In this paper we concentrate on a different example: a 1 +
1dimensional conformal field theory dual to a conical defect in AdS3 spacetime. We will
see that in this setting, the entwinement of certain fractionated degrees of freedom can
be understood in geometric terms by going to the covering space and applying the
RyuTakayanagi proposal. The upshot is that the entwinement of the fractionated degrees
of freedom corresponds in the dual conical defect spacetime to nonminimal geodesics
curves, which are local but not global minima of the distance function. This analysis
comprises section 2.
In section 3, we apply these results to the formalism we developed in [5], which
reconstructs analytically the bulk geometry from field theory data. We find that the conical
defect spacetime contains a macroscopically large entanglement shadow a central zone
surrounding the conical defect that cannot be reconstructed from entanglement entropies of
spatial boundary regions. The geometry in the central zone is determined by entwinement,
as we explicitly demonstrate. In fact, we will show that even outside the entanglement
shadow there are geometric objects whose boundary description involves entwinement.
The conical defect spacetime is similar to a black hole in having a macroscopic
entanglement shadow whose size is controlled by the mass of the object. In section 4 we discuss
how our findings inform the debate about reconstructing the geometry near and beyond
the horizon of a black hole.
Conical defects, long geodesics and entwinement
The entanglement shadow of the conical defect geometries
We begin with a review of the conical defect geometry. Starting from the global AdS3
ds2 = 1 +
R2 1
we obtain AdS3/Zn simply by declaring the angular coordinate to be periodic with period
r = R/n
t = nT ,
Spatial geodesics in the conical defect geometry descend from geodesics in the covering space with
one endpoint ranging over the n images. All but one of them are long geodesics.
ds2 =
2 1
means that the spacetime is singular on this locus.
Using the coordinate transformation (2.2), it is trivial to find the
spatial geodesics in the conical defect spacetime. One starts with the n geodesics in the
AdS geometry (2.1), which have one endpoint in common while the other ranges over
endpoints in the conical defect geometry. This is illustrated in figure 2. The geodesics are
described by the equations
and their length is:
= 2L log
size of the boundary interval subsumed by the geodesic, including winding. For example, a
geodesic that winds twice completely around the conical defect and returns to the original
the infinite tails of the geodesics near the spacetime boundary.5
geometry describes a pure state. Even if there is a mixed state with the same geometric
description, it would not have macroscopic entropy and the associated horizon would
have to hug the conical defect, and hence the RT prescription, which in this case includes
the horizon area, would still give the same answer.
rcrit(n) =
Thus, the central zone r < rcrit(n) is not probed at all by entanglement entropies of spatial
regions of the boundary. We call this zone of spacetime the entanglement shadow.
the interpretation of computing entanglement entropy, because they do not satisfy the
out to compute entwinement the novel concept that is the subject of the present paper.
In section 3 we will show how the full conical defect geometry, including the entanglement
shadow, can be reconstructed in the formalism of [5] using boundary entwinement as input.
Ungauging the dual description of the conical defect
There are various ways of representing the conical defect spacetime in a dual field theory.
One approach is to regard the defect as an excited state of AdS3. In this picture, we
scale and G is the Newton constant in three dimensions. The vacuum of this theory is
empty AdS space and the conical defect is a particular excited state. We will discuss this
description further below, but turn at present to a more convenient view of the system in
terms of its covering space.
5If we can compare cutoffs by matching radial positions in a standard FeffermanGraham expansion near
infinity, then perhaps one should rescale the cutoff with a factor of n as well. Such a rescaling would yield
a simple additive contribution to the entanglement entropy which we will ignore in the remainder.
As described above, AdS3/Zn can be regarded as an angular identification of a covering
AdS3 spacetime. The covering space ungauges the Zn discrete gauge symmetry and
physical quantities are computed by considering Zninvariant quantitites in the ungauged
theory. Indeed, the correlation functions of quantum fields in the conical defect and BTZ
spacetimes are typically computed precisely by taking this sort of view, which is equivalent
to the method of images for computing Green functions [43, 44]. Boundary limits of these
Green functions correctly compute the correlation functions of the corresponding CFT
states [44, 45]. This does not say that AdS3 with a conical defect is exactly identical to
the covering theory in its ground state; in fact it is not. However, many Zn invariant
observables and correlation functions computed in the covering theory agree with their
corresponding counterparts in the conical defect theory.
From this perspective, the field theory dual to the conical defect should also be lifted
to the covering space, which is an ntimes longer circle. We will denote this parent theory
CFTc, where c is a new central charge to be determined later. Spatial locations x in CFTc
lift to locations x in a fundamental domain of the covering space and to the corresponding
Zn translates. Correlations functions of CFTc that descend to CFTc must be Znsymmetric
ator and g is a Zn generator. This recapitulates the method of images used in the bulk to
compute the same correlation functions [44]. In the geodesic approximation [45] the
correlation functions between O1(x) and O2(y) would be computed from the geodesics between
lifts (x, y) of (x, y) to the covering space, and between all Zn translations of these locations.
The geodesic between x and y in the covering space descends to the minimal geodesic on
the defect, and the geodesics between x and the Zn translates of y descend to the long
geodesics on the defect geometry. The leading contribution to the correlator comes from
the minimal geodesics in figure 2, but the long geodesics yield subleading saddle points and
are all necessary to give the correct correlation function in the defect theory.
What is the central charge c of the covering CFT? We will give three arguments
algebra of the AdS3 spacetimes [46]. In this construction the central charge c of the covering
space is derived from the Virasoro algebra of large diffeomorphisms of spacetime:
[Lk, Ls] = (s k)Lk+s +
But not all such diffeomorphisms of the covering space will descend to the defect theory.
To preserve Zn symmetry we must restrict to a subalgebra generated by Lnk, because Lk
are Fourier modes of the boundary deformations. Following [47] we recognize that this
subalgebra also has Virasoro form with the following redefinition of the generators:
Lk =
k 6= 0
L0 =
L0 24
This is the Virasoro algebra of deformations that descend to the defect theory; it has a
3L/2G, we will interpret this relation as also saying that the covering theory has a rescaled
string picture (right). The short string interval maps to a union of disjoint long string intervals.
The geometries on the left represent the spatial slice of the conical defect while the geometries on
the right are spatial slices of antide Sitter space, which is the nfold cover of the conical defect.
Here n = 5.
This indicates that the CFTc dual to the defect theory has a fractionated spectrum, where
momenta are quantized in units of 1/n times the length of the spatial circle. Below we will
see direct evidence for this in a weak coupling limit of the CFT, where it can be visualized
in terms of an nwound string.
To test our identification of c let us compare entanglement entropies computed directly
entanglement entropy in the large c limit is [48, 49]:
+ const. =
The second equality is the RyuTakayanagi proposal (eq. (1.1)); it follows from eq. (2.5)
S(R) =
The relevant geodesics in the conical defect spacetime are displayed in the left panel of
The interval R in CFTc lifts to n evenly spaced intervals R i, each of angular size
S(R) =
entropy of this union of intervals is computed from the length of the minimal curve in
empty AdS3, which is homologous to their union. As shown in the right panel of figure 3,
+ const. = S(R) ,
carried out in [50].6 The relevant geodesics appear in families of n identical images, which
guarantees Zn invariance of the result. Here, Zninvariance is obtained, because the input
symmetrized. From the perspective of the covering theory, the transition between the
geodesics that subtend 2/n or 2( )/n marks two phases where disjoint intervals R i
and R i+1 do or do not share mutual information [51]. Interestingly, this mutual information
transition in the CFTc conspires to correctly reproduce the entanglement computations for
a single interval in the defect theory, where there is no mutual information to account for.
The weaklycoupled limit of the CFT dual to the conical defect is also very instructive
concerning the above points. To arrive at these insights, we first give a lightning review
of the relevant facts about the duality relating asymptotically AdS3 spacetimes and the
D1D5 field theory. Consider N1 D1branes wrapped on an S1 and N5 D5branes wrapped
T 4. The low energy description of this system is a 2d CFT, whose moduli space
horizon limit of the geometry sourced by this Dbrane system is AdS3 S3
the AdS curvature scale L proportional to N in threedimensional Planck units. A weakly
coupled type IIB string theory on this geometry is dual to a certain marginal deformation
of the orbifold CFT with large N . The weakly coupled limit of the CFT is near the orbifold
point and corresponds to a strongly coupled AdS theory. We are going to consider this
limit. The lowenergy CFT describing the brane dynamics is identified with the theory
dual to AdS3, living on the conformal boundary of this space:
The central charge of the CFT is c = 6N .
We are interested in the geometry AdS3/Zn. Its conformal boundary is also (2.12), so
this geometry should be dual to a state in the D1D5 CFT. At the orbifold point this dual
has been identified [44, 5254] as the state
fields in the CFT are single valued on the nfold cover of a spatial slice of the theory. More
6Note that eq. (2.11) receives 1/c corrections while eq. (2.9) is exact. This is to be expected: CFTc in
state (2.13) is not identical to the ground state of CFTc.
Figure 4. A strand of three target space fields X1,2,3, which define a single field X 1 of the long
explicitly, a field configuration in the twisted sector is given by the profiles of N T 4valued
target space fields around the worldvolume circle, which we call X1, . . . , XN . The twisted
around the S1, the fields transform into one another as N/n strands, each containing n
XNn+1
XNn+2
XNn+1
This means that we can equivalently represent the theory with N/n singlevalued fields
on a circle, whose circumference is n times longer than the circle supporting the orbifold
by gluing together the values of the fields in (2.14), as illustrated in figure 4. They define
the worldvolume theory of the long string [30].7 So long as we consider untwisted probes
and excitations, the dynamics is restricted to the superselection sector of this long string.
The construction of the long string on the nfold cover of the short string means that we
can think of the long string as living on the boundary of the covering space of the defect,
which, as we discussed above, is simply empty AdS3 space. Imposing the BrownHenneaux
relation [46], we think of this AdS3 cover as having a rescaled Planck constant G [47].
Because the long string is n times longer than the short string, momentum on its
the spacings of the energy levels. The reduction in energy gap is called fractionation.
This reproduces our observation above based on the Virasoro algebras of the CFTc and
CFTc. The factor n also relates the central charges of the two theories: encapsulating the
fields X1, . . . , Xn in a single field X 1 trades n degrees of freedom for a single degree of
7According to the standard orbifold prescription, we should still mod out the Hilbert space of the N/n
long strings by the the action of Zn on each of the short strings, and also by the action of the permutation
group SN/n which exchanges the long strings among themselves. This is consistent with the covering CFT
still being an orbifold theory, but now one based on SN/n.
and c that we derived above from symmetry considerations and verified using entanglement
came from the state (2.13) of CFTc through the restriction on its set of gauge invariant
observables and the action of Zn on its Hilbert space.
Entwinement and entanglement shadows
In section 2.1 we saw that the conical defects have an entanglement shadow a central
region which is not probed by CFT entanglement. This shadow exists because the minimal
geodesics in the RT formula for holographic entanglement only penetrate to a certain
the defects or wind around it) do penetrate the entanglement shadow, but they do not
contribute to the entanglement entropy, at least at the leading order, according to the RT
formula. Nevertheless, the long geodesics are certainly related to physical quantities in
the CFT. As discussed in section 2.2, they make subleading contributions to boundary
correlation functions, and are in fact necessary for conformal invariance. Thus, we may
wonder whether the long geodesics should also make a subleading contribution to the
entanglement entropy (thereby modifying the entanglement shadow), perhaps via simple
additive pieces resembling their methodofimages contributions to semiclassical correlation
functions. Such a picture is too simplistic if, as discussed above, the conical defect can be
regarded as a pure excited state of the D1D5 string. In this case, as the CFT interval
tends to the size of the entire boundary, the entanglement entropy must tend to zero, which
it will not if we include contributions from the long geodesics (e.g. from the geodesics that
start at a point, wind around the defect, and return to the same point).
The covering space picture in figure 3 further illuminates the problem. As discussed in
size (2 2)/n between the Ri, precisely when the intervals Ri begin to share mutual
information. If not for this mutual information transition in the covering space theory,
the geodesics subtending the Ri would have descended to long geodesics in the conical
covers the entire boundary and there is no entanglement entropy to be considered since
the state is pure. By contrast, single boundary intervals Ri in the covering space of size
geodesics on the conical defect.
We see that from the covering space perspective long geodesics are eliminated in the
holographic spatial entropy formulae by a mutual information transition that arises,
because spatial entanglement entropy in the conical defect is computed from the entanglement
of a Zninvariant union of intervals in the covering space theory, i.e.
where g is a Zn generator (see figure 3). Physical observables that descend to the
conical defect certainly must be Zninvariant. Are there Zninvariant quantities related to
entanglement in CFTc that can be computed without first taking the Zninvariant union
of intervals in that theory? One possibility is to compute the entanglement entropy of a
single interval and then sum the result over Zn translations:
E(R) =
X S R i =
n1
This quantity, which we call entwinement, is Zn invariant and thus descends to the conical
computed in the covering AdS3 by a minimal curve, which descends in the conical defect
to a long geodesic that penetrates the entanglement shadow (see figure 2). Explicitly, for
E(R) = n 4G log
We have used 3L/2G = c = c/n in a manner analogous to eq. (2.11).
Let us summarize the steps we have taken to define entwinement. We ungauged the
discrete Zn symmetry of the conical defect theory, computed conventional spatial
entanglement in the parent theory, and then symmetrized the computation to get a Zninvariant
quantity we called entwinement. How can we interpret this quantity directly within the
conical defect theory? Recall that we argued that CFTc, which is dual to the conical
defect, has a set of internal degrees of freedom with fractionated energies and momenta.
We propose that entwinement captures the entanglement of subsets of these degrees of
freedom in given spatial regions with the rest of the theory. This interpretation is
easiest to visualize in the longstring picture of the dual, which appears close to the orbifold
point. As described in section 2.2, in this picture there is an effective string with central
total of c degrees of freedom at each point. Entwinement computes the closest analog to
entanglement that applies to a partition of the windings into subsets. If we compute the
of one winding, summed over windings (i.e. summed over Zn translations). In the range
effective string, summed over Zn translations. Our procedure of removing the Zn
identifications and symmetrizing afterwards recalls methodologies that have been used before
to study conventional entanglement entropy in gauge theories (see [5558] and references
To emphasize the role of ungauging from a slightly different angle, consider a Hilbert
into themselves. This is the case if, for example, H1 is the Hilbert space associated to the
union of short intervals in the long string (whose associated geodesics descend to minimal
geodesics in the conical defect) and H2 is the Hilbert space associated to the complement.
Even though G preserves H1,2, this tensor factor decomposition does not descend to a
tensor factor decomposition of HG. We can certainly decompose H1 and H2 into irreps Ri
of G, so that the decomposition reads
G will only contain Gsinglets, and therefore only contains contributions from the
project in general on Ginvariant states. One is always left with a complicated sum of
tensor factors. Thus, even for a short interval, the appropriate notion of entropy cannot be
obtained as the entanglement entropy associated to a tensor factor in the invariant Hilbert
G and one always needs to ungauge.
For sufficiently long intervals, associated to long geodesics in the conical defect, there
preserves H1,2. What we have effectively done is to pass to an even bigger Hilbert space
Hextended = gGg(H1) g(H2)
whose decomposition is now compatible with an action of G (it permutes the summands).
It would clearly be interesting to explore the connection between such decompositions and
the discussion in [5558] (and references therein).
Reconstruction of geometry from entwinement
Geometry from entanglement
In [5], we showed how to compute the circumference of an arbitrary, piecewise differentiable
closed curve on a spatial slice of 2+1dimensional antide Sitter space from boundary data.
For definiteness, we work in global coordinates
ds2 = 1 +
R2 1
determined in one of two ways:
(1 ) The outgoing null ray orthogonal to the curve reaches the boundary after a global
Then the length of the closed curve is given by the formula:
1 Z 2
The colorcoded summands correspond to the continuously drawn pieces of the geodesics in the
figure. The difference between their lengths aligns with the length element along the central circle
R = R0.
The second equality uses the RyuTakayanagi relation (1.1). In the special case of a central
explains formula (3.2) from the bulk point of view. For a general bulk curve, a similar
explanation holds, though the technical details are more involved. For more information,
Reconstructing the conical defect spacetime
The conical defect spacetime is locally AdS3. Therefore, the middle formula in eq. (3.2),
which is an identity in the bulk, extends to the conical defect automatically. The version
on the right hand side, however, is a boundary statement, which applies only so long as
is when construction (2 ) in section 3.1 above returns a long geodesic, the set of spatial
boundary entanglement entropies is insufficient to define the given bulk curve, let alone to
calculate its length. As we saw in eq. (2.6), this occurs whenever the curve approaches the
regime, the entwinement computed in eq. (2.17) is a necessary ingredient.
In fact, long geodesics and entwinement are also necessary to recover the metric outside
this central zone. In particular, any closed curve with a sufficiently radial local direction
The long geodesic tangent to it is shown in green and the tangency point is marked with a circle.
shown in dashed red) is tangent to a long geodesic and therefore probes entwinement. In particular,
every curve that is locally parallel to the radial direction probes entwinement. In the entanglement
shadow (r < rcrit, marked in dashed purple) all curves probe entwinement, regardless of the slope.
A curve, which at any radial scale approaches the boundary more rapidly than
geodesic (3.4), cannot be defined or measured in the boundary using spatial entanglement
entropies alone. (See the related discussion in [6].) Importantly, without this information
one cannot determine the radial component of the bulk metric anywhere. The
entanglement shadow in the region r < rcrit is special in that there even the angular component
of the metric is inaccessible from boundary spatial entanglement entropies. The critical
geodesic (3.4) and its maximal depth rcrit(n) should be relevant to the problem of bulk
locality on subAdS scales; it would be interesting to flesh out this relation.
length (2.5) into the middle expression in eq. (3.2), we obtain:
1 Z 2
1 Z 2
given in eq. (2.17) in terms of entwinement. More generally, entwinement is necessary to
define and measure any curve with a sufficiently radial direction anywhere, see figure 6.
We have argued that entwinement an analog of entanglement for gauged Hilbert spaces
which do not contain all the degrees of freedom of spatial regions in a field theory is an
essential ingredient for holographically reconstructing the conical defect geometries in three
dimensions from field theory data. There is a central zone near the conical defect, which
is not probed by conventional spatial entanglement in the field theory and entwinement
becomes necessary. Furthermore, even far away from the defect, entwinement is implicated
in the emergence of the radial direction of space.
One of our lines of argument involved the orbifold description of the weakly coupled
D1D5 string, which made the discrete Zn gauge symmetry easy to visualize and manipulate.
The field theories that we encounter in AdS/CFT are usually low energy limits of gauge
theories, and we therefore expect that some notion of gauge invariance is still present away
from the weak coupling limit. This is illustrated directly by the CFT that is dual to the
covering space description of the conical defect, and which is the strongly coupled version of
the long string sector that is evident at weak coupling. Therefore, we expect the existence
of a fieldtheoretic definition of entwinement away from weak coupling as well.
The long geodesics associated to entwinement map to spatial intervals that cover
the boundary of the conical defect more than once, suggesting that the associated
observers have access to all the information available in the field theory. However, following
method (1 ) of section 3.1, one can show that the associated time intervals on the boundary
are finite. It was proposed in [5] that ignorance of the quantum state associated to such
finite time measurements (called residual entropy in [5]) is related to the areas of closed
curves in the bulk spacetime and their associated entropies [4]. Indeed, in the present
case, the observers associated to time intervals that are too short will not have the energy
resolution to probe excitations with the fractionated gaps of the conical defect. More
precisely, it is clear from the bulk point of view (employing the covering CFT), as well as from
the long string point of view, that by causality such observers cannot fully access all Zn
invariant correlation functions. Entwinement is a quantity which might be associated to
this ignorance and it would be very interesting to make the connection more precise. One
can perhaps get some clues from recent work by Hubeny [59], which investigates potential
covariant definitions of the residual entropy associated to finite time observers, and possible
relations (or the lack thereof) to minimal holographic surfaces.
The role and robustness of entwinement would be clarified by a path integral definition
of entwinement, perhaps starting from a notion of Renyientwinement. Likewise, it would
be interesting to study the role of entwinement in higher spin theories, where the notions of
conical defects and covering spaces can still be defined in terms of monodromies of gauge
fields, and most of our discussion naively still seems to apply. Even more generally, if
entwinement is indeed related to fractionated degrees of freedom, one would expect it to
be of relevance in all theories with a HawkingPage like phase transition at large central
charge. We leave these questions as open.
We believe that these lessons are not limited to conical defect geometries, but apply in
greater generality. Consider, for example, the massless BTZ geometry [60] whose metric
whenever we discuss geometric quantities in the bulk. Thus, the results of the present
particular, the massless BTZ spacetime contains spatial geodesics that wrap around the
black hole infinitely many times, as opposed to the maximal number of n/2 for AdS3/Zn
(see eq. (2.5)). The entanglement shadow, which short geodesics do not reach, extends
The same critical radial scale can be found directly from the form of spatial geodesics
rcrit() = lim rcrit(n) =
n
r2 = r+2 sinh2(r+/L) sinh2(r+/L)
in the massive, stationary BTZ metric
ds2 =
r2 r+2
Beyond the massless BTZ geometry, a qualitatively new ingredient appears. When we
obtain the conical defect geometry as an orbifold of AdS3, we mod out by a finite subgroup
of rotations. Formally speaking, we orbifold by an elliptic element of the conformal group.
The massive BTZ geometries are obtained from orbifolding by a hyperbolic element, in
other words, a boost.9 The argument presented in this paper identified the boundary of
the AdS cover with the worldvolume theory of the long string. In the BTZ case, this
identification is more complex. First, the orbit of a boost is noncompact, so the covering
space contains infinitely many copies of the BTZ spacetime. Second, the boost acts
noncomplications is likely to teach us more about the emergence of horizons from entanglement
in the dual boundary theory.10
tanglement plateau [20] are also short in the sense that they compute entanglement entropies of spatial
regions on the boundary. This is due to the homology constraint in the RyuTakayanagi proposal. But in
from the ArakiLieb inequality [61]:
S(R) S(Rc) S(R Rc) = SBTZ = 0 .
9The massless BTZ is the critical case of orbifolding by a parabolic element. As we saw in the previous
paragraph, it can be recovered either as a limit of elliptic or hyperbolic orbifolds.
10Note, however, that several authors [6266] have suggested subtleties that might challenge the
interpretation of the BTZ black hole as an orbifold at the quantum level.
For the weakly coupled orbifold theory, we can extend our proposal to more general
states and in particular thermal states. The full Hilbert space then involves a sum over
tensor products of strings of various lengths, and by summing the entanglement of long
intervals over those the twisted sectors that can accomodate long intervals with appropriate
weights, we obtain a natural generalization. It would be interesting to do this computation
for a thermal state and to compare to the lengths of long geodesics (i.e. nonminimal
geodesics that wind around the horizon) in the massive BTZ background.
Spatial entanglement, entwinement and mutual information
As formulated above for the case of the conical defect, the difference between spatial
entanglement and entwinement lay in the order in which we carried out the Zn symmetrization:
This difference can be traced to the nonvanishing mutual information between the image
highlighted in (4.5):
I(R 1 : R 2) = S(R 1) + S(R 2) S(R 1 R 2)
At least in the special case of AdS3/Z2, the mutual information between images can
therefore be thought of as an order parameter for the relevance of entwinement in a holographic
reconstruction of spacetime.
Toward higher dimensions
In the conical defect geometry entwinement computes areas of extremal but nonminimal
curves. In higher dimensions, extremal but nonminimal surfaces likewise play a role in the
emergence of holographic spacetimes. One example was given in [7, 8], which generalized
formula (3.2) to compute areas of codimension2 surfaces in higherdimensional spaces. Like
eq. (3.5), that computation generally involves extremal but nonminimal surfaces. It would
be fascinating to lift the definition of entwinement, which in its present form pertains to
2d CFTs, to give a boundary interpretation of areas of extremal but nonminimal surfaces
in higher dimensions. Likewise, a generalization to higher curvature theories would be
illuminating, especially in light of the presence of closed extremal surfaces [67].
We saw in section 3.2 that entwinement is relevant in boundary computations of bulk
lengths under two distinct circumstances: when the bulk curve becomes nearly radial
and when it probes the region near the conical singularity. The computations in [7, 8]
are roughly analogous to the former circumstance, but does the latter have an analogue
in higher dimensions? More specifically, when do higherdimensional spacetimes possess
entanglement shadows regions outside the reach of minimal surfaces?11 Ref. [20] showed
that AdSSchwarzschild black holes are similar in this respect to BTZ black holes: they have
11We thank Mukund Rangamani and Mark Van Raamsdonk for clarifying this issue and for pointing out
entanglement shadows that surround the horizon up to a thickness of order LAdS. On the
other hand, it is known that horizonless geometries with matter can have no entanglement
We are grateful to Michal Heller for his collaboration in the early part of this project and
for many discussions. We thank Dionysios Anninos, Xi Dong, Liam Fitzpatrick, Patrick
Hayden, Matthew Headrick, Jared Kaplan, Nima Lashkari, Javier Magan, Rob Myers,
Mukund Rangamani, Joan Simon, James Sully, Mark Van Raamsdonk, Erik Verlinde, and
Herman Verlinde for helpful discussions. VB was supported by DOE grant
DEFG0205ER41367 and by the Fondation PierreGilles de Gennes. The work of BC has been
supported in part by the Stanford Institute for Theoretical Physics. This work is part of
the research programme of the Foundation for Fundamental Research on Matter (FOM),
which is part of the Netherlands Organisation for Scientific Research (NWO). This work
was supported in part by the National Science Foundation under Grant No. PHYS1066293
and the hospitality of the Aspen Center for Physics. BC dedicates this paper, submitted
on Fathers Day, to his Dad, Stanislaw Czech.
This article is distributed under the terms of the Creative Commons
Attribution License (CCBY 4.0), which permits any use, distribution and reproduction in
any medium, provided the original author(s) and source are credited.
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