# Measurements in general relativity

[THINGS TO FIND : Dual foliation, can any (finite-acceleration) curve fit a foliation (cf acceleration as derivative of ln(N) for that maybe), is the exponential map bijective in some tubular neighbourhood of an observer, find net of synchronizations]

A physical theory is more than simply its mathematical structure. It also requires some mapping of those elements to real life measurements. Although such a mapping is always difficult, or at least somewhat arbitrary, this is trickier than it appears in general relativity due to our lack of knowledge of the geometry of the overall spacetime. Many experiments in general relativity have some fairly broad assumptions to allow us to assume the spacetime metric, but is it actually possible to measure it with a minimal amount of assumptions?

As general relativity is a theory of spacetime, our basic measurements to consider will be distances, angles and durations. There are more quantities we could measure of course, but these will be the basic notions we'll need for informations on our metric.

As the goal is the experimental measurement of a spacetime, things will be done in two steps here : First, given a known spacetime, try to find if this gives rise to unique measurements, and then, given a set of such measurements, try to find out which spacetime we are dealing with here.

While the attempt will try its best to remain broad, for our sanity, a few assumptions will remain here :

- Neither atomic clocks nor rigid rods are impervious to the spacetime metric nor the elements, but we will assume that those indeed define the measurements we want. At most, we'll allow for some known bounded uncertainty $\Delta t$ and $\Delta d$

There are of course many types of experiments we can do, but we'll consider the following scheme here. These are fairly idealized ideas, but we'll see later on what will happen when we decide for a more realistic set of measurements.

Through every point of spacetime, an observer, idealized by a timelike curve, passes through. Each such observer is identified by some indexes (for instance if our spacetime has the Cauchy surface $\mathbb{R}^3$, then the index could be $(x,y,z)$). Each observer has on board a clock and accelerometer, and it can fire off either light or particle beams of arbitrary energy. Each such beam is encoded as to contain the relevant informations : when it was fired, according to the local clock, at what angle it was fired, and the index of the observer firing it. Those beams can be received by other observers, but they will also continue on beyond that point (for a more realistic perspective, we can say that the observer intercepting it will simply relay it by firing it again, perhaps adding its own informations to it), and it will fire back to the emitter of that signal with its own informations.

In other words, the emitter will emit a signal of the form

\begin{equation} \text{Emitted at $t = T$ by $(x,y,z)$ at angle $\theta, \phi$} \end{equation}And may receive something of the form

\begin{equation} \text{Signal originally Emitted at $t = T$ by $(x,y,z)$ at angle $\theta, \phi$, received at $t_R = T_R$ by observer $(x_R, y_R, z_R)$} \end{equation}There are many reasons why this idealization cannot work : we obviously cannot have one such machine per point of space, idealizing such a complex machine as a single point is quite a lot, sending so many signals in infinitely many directions will both raise issues with the quantity of energy that you would need to store for such a thing as well as break the idealization of those observers in the test field limit, the quantity of data to transmit and store would be literally infinite at the precision we're requesting (both the clock value and the index which is part of $\mathbb{R}^3$, for a start), and of course all the quantities we're using have some inherent uncertainty from real measuring apparatus or simply the process of emission (sending a light wave with so much data will certainl). Once we have the theory for the exact measurement using this idealization, we'll try our best, with minimal assumptions, to find out what a more realistic case looks like.

The more well-known case for all of this is of course Minkowski space, so a lot of our first examples will be (for simplicity) two-dimensional Minkowski space, in other words $(\mathbb{R}^2, \eta)$ with the canonical coordinates $(t,x)$ such that

\begin{equation} \eta = -dt \otimes dt + dx \otimes dx \end{equation}As all our measurements are done with causal means, and that we probably cannot assume an experiment spanning the whole duration of our universe, what we essentially get is a measurement of spacetime equivalent to a causal diamond. If our main observer starts at $p_1$ and ends at $p_2$, the spacetime probed will be fundamentally $I^+(p_1) \cap I^-(p_2)$. In other words, we won't be able to probe anything outside the horizon of our observer, as is to be expected. We can hopefully perform some spacetime extension from our data that could be unique enough to have a reasonable idea of what it may be.

## 1. Measuring time in general relativity

### Clocks

Asserting the timing of events in experimental general relativity is the first thing we need to do, for all other quantities will stem from the foliation of our spacetime into spacelike hypersurfaces, so that we can use proper Riemannian geometry to compute the distances and angles.

While there are certainly quantities related to time for theoretical entities in general relativity, these are fairly arbitrary, although this is also true of classical theories, where the point of origin, direction and scale of time measurements are somewhat arbitrary. There's no way to provide measurements of the time coordinate from the manifold's atlas, nor from any time function we may define, as we can change any such quantities via diffeomorphisms or a different slicing of our manifold. We can't even guarantee that such a foliation of our spacetime into spacelike hypersurfaces exist, so that our slicing even makes sense.

As we (probably) cannot get global informations from our spacetime manifold, all definition will therefore have to be local. Therefore, we'll require all our definitions to be associated with an observer $\gamma$, represented by a timelike curve. An obvious number we can associate with the observer is the parameter $\tau$, such that we can define our time for the observer as

\begin{equation} \forall p \in \text{Im}(\gamma),\ t(p) = \tau \text{ such that } \gamma(\tau) = p \end{equation}but of course this parameter is fairly arbitrary, as we can simply redefine a parameter for our curve. Therefore, we have to tie our definition of time to a measurable process.

If we look at modern physical theories, there doesn't seem to be any process which is exactly periodic, but this does not matter. Our notion of time is defined independently of any such considerations. We'll therefore define a notion of clock here :

**Definition :** A *clock* in general relativity is a process which can be approximated by a timelike curve for which there are events separated in (measured) time $t$ by congruent durations. In other words, there exists a duration $T \in \mathbb{R}$ such that for every events $e_k \in \gamma$, $k \in \mathbb{Z}$,

In other words, every such occurences of the event is defined to be $T$ of our measuring unit from the previous one. If we define some original event $e_0$ such that $t(e_0) = 0$, then we end up with the measurement of time

\begin{equation} t(e_{k}) = k T \end{equation}or more generally $t_0 + kT$ if we define $t(e_0) = t_0$.

#### Atomic clocks

A particularly useful standard clock is the one given for the second.

**Definition :** A *second* is the interval of time between $9 192 631 770$ maximal amplitudes of the radiation obtained by transition of the two hyperfine levels of the ground state of Cesium 133 at rest and at $0\ K$.

Therefore, if we have an atomic clock along with our observer, we'll define a metric (as in metric system here, not metric tensor) parametrization of the curve if, for some initial time parameter $\tau_0$, and given the original proper time parametrization $\tau$ and the interval between two maximal amplitudes as $T$, our new parametrization is

\begin{equation} \tau_m = \frac{\tau - \tau_0}{T} \end{equation}Unfortunately, all we have so far is a way to measure only the observer's own time, which is of limited use. The next step is to find a way to assign a time to non-local events.

#### Light clocks

Another common way of measuring time in general relativity is the notion of light clocks.

#### Influence of gravity on the clocks

No matter what happens, if our spacetime is globally hyperbolic anyway, our clocks will define some subset of a time parametrization for our observer, ie, $\gamma(t_0 + k T)$ is such that there exists a parametrization where this is the correct measure of time. What we would like in the ideal case as well is that the clock also measures the proper time, as this would simplify a lot of our issues. It's quite easy to see that for instance, if our clock only ticked for every $\tau^3$

All of our clocks fundamentally rely on electromagnetism and the Dirac equation to work : either via the transition frequency of electron in an atom or via the length of a rod, which depends roughly on the atomic radius. But the Dirac equation is something of the form

\begin{equation} i \gamma^a e^\mu_a (\partial_\mu - \frac{i}{4} \omega^{ab}_\mu) \sigma_{ab} - i Ze A^\mu) \psi - m \psi = 0 \end{equation}Which leads us to the question : does the influence of gravity on our measuring apparatus works out in the end, or does it induce some fundamental uncertainty? It's entirely possible that an atomic clock will still measure the proper time, but that is certainly not a trivial statement.

There are two main drawbacks to all of this. The first one being its discrete nature (we can't really get much better than the time between two crests), and it is itself sensible to the spacetime geometry (among other things, as it will also be sensible to a variety of other circumstances). Indeed, let's consider what happens here. Without getting too deeply into quantum field theory, let's consider some Riemann normal coordinates around the nucleus, which is considered fixed and classical, and look at the Dirac equation in curved spacetime (the Einstein-Maxwell-Dirac equation) for an electron :

\begin{eqnarray} e^\mu_a &=& \delta^\mu_a + \\ \omega^{ab}_\mu &=& \end{eqnarray}### Clocks and synchronization in special relativity

The simplest case we can work with here is of course special relativity, in Minkowski space. We'll consider the two dimensional case for simplicity, although this generalizes quite readily. In this case, let's consider two observers, $\gamma_1$ and $\gamma_2$. There's a few cases we should be mindful for.

Out of all of these, the simplest case is to consider two timelike geodesic curves. In Minkowski space, this is simply

\begin{equation} x_i(\tau) = (\gamma_i \tau + t_0, \beta_i \gamma_i \tau + x_0) \end{equation}which indeeds obeys the geodesic equation, and is timelike if

\begin{equation} \gamma_i = \frac{1}{\sqrt{1 - \beta_i^2}} \end{equation}our usual Lorentz factors. Also of use is the parametrization of a null geodesic,

\begin{equation} x_i(\tau) = (\tau + t_0, \pm \tau + x_0) \end{equation}If we perform the synchronization by exchange of light signals, $\gamma_1$ emitting $\ell_1$ at $\tau_1$, $\gamma_2$ receiving it at $\tau_2'$ before emitting $\ell_2$ immediately which $\gamma_1$ receives at $\tau_1''$, then the Reichenbach synchronization is, for $\varepsilon \in (0,1)$, < /p> \begin{equation} \tau_1' = \tau_1 + \varepsilon (\tau_1'' - \tau_1) \end{equation}

If we consider an observer $\gamma$ starting at $\gamma(0) = (0,0)$ emitting light signals in both directions at every instant $t$, $\ell^\pm_t$, we have that \begin{equation} \ell^\pm_t(\lambda) = (\tau + t_0, \pm \tau + x_0) \end{equation} any event $p = (t, x)$ will intersect $\ell^\pm_t$ as

\begin{equation} \tau_1' = \tau_1 + \varepsilon (\tau_1'' - \tau_1) \end{equation}The more common synchronization is of course Einstein's synchronization, which is simply $\varepsilon = 1/2$

\begin{equation} \tau_1' = \tau_1 + \frac{1}{2} (\tau_1'' - \tau_1) \end{equation}### Synchronization in general relativity

Things get much more complicated in general relativity due to the fact that we do not know the intermediate geometry. There are many things which can make clock synchronization painful here.

First, the local time at which we receive the signal back may happen before we've sent it, for instance by taking the Deutsch-Politzer spacetime :

A signal sent may never come back at all, for a variety of reasons. If we attempt to send a signal from $\gamma$ to $q$, $p_1 \in \gamma$ the time of signal, perhaps $q$ does not belong to $I^+(p_1)$, or $\gamma \cap I^+(q) = \varnothing$, or there may simply be no null geodesics connecting the two (consider Minkowski space with a point removed as an example).

A global synchronization may simply not exist. Even under the broadest definition of a synchronization (perhaps just assuming some foliation by spacelike hypersurfaces), some spacetimes, such as the Gödel spacetime, will not admit such a thing.

There may be some discontinuities in the light clock signals if the topology isn't trivial. Consider a wormhole spacetime with an outside observer. With our plan to send a signal at every point of our observer's curve, what we are basically getting if we retract our entire spacetime along its foliation is a family of loops, which will locally be contractible to the identity. As the fundamental group of a wormhole spacelike hypersurface isn't trivial, we will finally get a curve that cannot be contracted to the identity, and two light signals $\ell_{t}$, $\ell_{t + dt} can then arrive at arbitrarily far apart times.

To simplify things somewhat, first, let's assume that $(M, g)$ is globally hyperbolic. While not necessarily uncontroversial, this will help out immensely to reduce the number of pathological cases to consider, and it will be at least a valid assumption in some neighbourhood (although how far from it is of course matter to some debate).

We now get a rather useful set of theorems :

- There exists a set of foliations of $M$ into Cauchy surfaces (achronal spacelike hypersurfaces) $\Sigma_t$, $M \cong \mathbb{R} \times \Sigma$, such that $\Sigma_t \cong \Sigma_s$ for every parameter $s,t$, such that every observer $\gamma$ crosses every Cauchy surface exactly once.
- Every foliation into Cauchy surface induces a parametrization on every observer $\gamma$ by $\forall \tau \in \mathbb{R},\ \gamma(\tau) \in \mathbb{R}^3_\tau $
- Globally hyperbolic spacetimes admit continuous time functions
- For two causally related points $p \leq q$, there is a non-spacelike geodesics of maximal length
- There is a foliation of $M$ by timelike geodesics
- The spacetime is isometric to the product spacetime $\mathbb{R} \times \Sigma$ with the metric $$ds^2 = -\beta(t, x) dt^2 + g_{ab}(x,t) dx^a dx^b$$with $g_{ab}$ a Riemannian metric on $\Sigma_t$. $t$ defines a time function on this spacetime and $dt$ is a tangent $1$-form to it.

#### The set of null geodesics

A very important subset of our spacetime that we will require is the set of null geodesics. For this, we define by $N^\pm(p)$ the set of all points $q$ such that there exists a null geodesic connecting $p$ and $q$. As most of our measurements will be done using null geodesics, it will be a useful set to have.

Unfortunately, the set of null geodesics isn't quite as trivial as we would hope. In other words, $N^\pm(p)$ is not actually equivalent (in general) to the horismos, $H^\pm(p) = J^\pm(p) \setminus I^\pm(p)$. A very simple example for this is to consider the Minkowski cylinder $\mathbb{R} \times S$, where null geodesics are helixes spiraling around but for every $2\pi$ turns, any two such points on the helix can be obviously joined by a timelike curve.

**Theorem :** In a globally hyperbolic spacetime, for a timelike curve $\gamma$ using a light clock at $p_1$, receiving it at $p_2$, the point $q$, $q \not\in \gamma$, on which it is reflected is simultaneous (no causal curves link the two) to some point $p \in \gamma$, $p_1 \ll p \ll p_2$.

**Proof :** There is a future-directed null geodesic $\ell_1$ from $p_1$ to $q$ and a future-directed null geodesic $\ell_2$ from $q$ to $p_2$. Reversing the parametrization, this means a past-directed null geodesic $\bar{\ell}_1$ from $q$ to $p_1$. Therefore, $p_1 \in J^-(q)$ and $p_2 \in J^+(q)$. As our spacetime is globally hyperbolic, $J^-(q) \cap J^+(q) = \{ q \}$. Since our timelike curve must be connected, the only way for it to be in both the future and past light cone of $q$ would be to pass via $q$, which is not in the curve. Therefore there is a section of $\gamma$ such that every point in that section is simultaneous to $q$.

We therefore have a point in between those two points on the light clock with which we can synchronize that event. Unfortunately, $p_1$ and $p_2$ are not necessarily on the light cone itself. Two counterexamples of $p \neq \partial J^+(q)$ despite a null geodesic between $p$ and $q$ are simply the Minkowski cylinder $\mathbb{R} \times S$, where a null geodesic will obviously lie in $I^+(q)$ after a single turn, and for a topologically trivial example, ultra-compact object spacetimes, in which the presence of a lightring makes it that any observer stationary on this lightring will be

For now, let's proceed backward : let's assume that we have a foliation $\mathfrak{t} : M \to \mathbb{R}$, such that $\mathfrak{t}^{-1}(t) = \Sigma_t$. Our goal here will be to find out if there exists a synchronization convention such that, using some manner of synchronization method, we can reproduce this foliation without using previous knowledge of our metric.

By definition of a time function in a globally hyperbolic spacetime, our time function $\mathfrak{t}$ is a surjective function that grows monotonously along every future-oriented timelike curve. At minima, we'd like such a function to at least agree with our main observer $\gamma_0$ : for every point $p \in \gamma_0$, $\mathfrak{t}(p)$ has the same value as $\gamma_0$'s onboard clock.

Let's now work backward for a bit : consider an existing time foliation of our spacetime by a time function $\mathfrak{t}$, with a preferred observer $\gamma_0$. For every point $q \in M$, we define a light trip by

\begin{equation} \ell_q = \ell_{p_1 \to q} \cup \ell_{q \to p_2} \end{equation}Unfortunately, such trips may not be unique. What we will ask here, for now (this definition will not work in the most general of cases) is that

#### The simultaneity bundle

The situation that we have here may bring back memories of bundles here. We have a spacetime that is locally (and in this case, globally) diffeomorphic to a product, $M \cong \mathbb{R} \times \Sigma$, and for any point $p \in M$, we could map it to either $\mathbb{R}$ or $\Sigma$ : either $\pi_{t}(p) = t(p)$, where we have the time function $t$ playing the role of the bundle projection onto $\mathbb{R}$, or $\pi_{x}(p)$ is the position of our point on the spacelike hypersurface. A way to achieve this is to pick both a time function $t$ and a foliation by timelike geodesics $\gamma_x$, such that if $t(p) = 0$, $\gamma_x(p) = \iota(x, 0)$, and then the projection will be

\begin{equation} \pi_x(p) = \{ y | p \in \gamma_y \} \end{equation}Our simultaneity bundle is a bundle $M \cong \mathbb{R} \times \Sigma$, of base space $\Sigma$ and typical fiber $\mathbb{R}$. Every point $x \in \Sigma$ has a corresponding fiber $\mathbb{R}_x$ representing an observer in our spacetime. Things will get more complex if we assume our main observer to be accelerated, so for now let's only consider a geodesic observer.

Our geodesics and foliation here allows us to relate points of distant fibers. In other words, they are a form of *connection*. To see this, let's try to consider thing in a more gauge approach. As our typical fiber is $\mathbb{R}$ here, we can try to promote it to a principal $\mathbb{R}_+$-bundle by considering the translation group. Our right-action on the bundle is

So that the bundle action is simply to travel along our geodesic foliation further in time by $T$. This action indeed preserves the fiber.

We'll consider the effect of our synchronization by observing it in the typical fiber obtained by quotienting out the bundle group, ie

\begin{equation} \Sigma = M / G \end{equation}In this case, any light-clock is simply a loop in $\Sigma$ with base point $\gamma(0)$. Unlike the case of Migunzzi[6], our spacetime isn't static, meaning that we cannot assign any induced metric here. We can't assign our synchronization simply given $\Sigma$, we have to work with every slice $\Sigma_t$.

The benefit here is that the tangent bundle of our bundle is simply the tangent bundle of our spacetime, ie

\begin{equation} TE = TM \end{equation}As we'll

The clock synchronization can be modelled after holonomies : given a point $p$ for our observer,

From all these elements, let's consider a fairly simple case : the class of all conformally Minkowski two-dimensional spacetimes. We're making this choice as it is both simple, and by Kulkarni's theorem, covers most spacetimes with $\mathbb{R}^2$ topology under some fairly reasonable assumptions. We therefore have to find out a nowhere-vanishing function $\Omega$ such that

\begin{equation} ds^2 = \Omega^2(x,t) (-dt^2 + dx^2) \end{equation}#### ???

As we are working with infinitely many signals and infinite precision, it remains possible here to consider the convex normal neighbourhood without having to worry about its extent. We just need to first determine what behaviour we will see once leaving it.

For our preferred observer, we can consider the neighbourhood defined by, for some time interval $I = (\tau, \tau + \Delta \tau)$, its tubular neighbourhood. For every $\tau \in I$, the exponential map has a certain injectivity radius $\rho_\tau$. If we take the radius

\begin{equation} \rho = \min_{\tau \in I} \rho_\tau \end{equation}Then all of our neighbourhood is appropriately normal. In our entire neighbourhood, we can simply map behaviour from the tangent space to the manifold itself.

## 2. Measuring distances in general relativity

Despite a few issues, measuring time in general relativity remains simple enough to work out with minimal assumptions. Distances, on the other hand, being rather non-local, are much more difficult to establish.

The two fundamental process for establishing distances in general relativity are *rulers* and *light signals*. Rulers are, generally speaking, objects that we suppose to be of fixed length, while light signals are done by an exchange of light signals between two objects, as we can assume such signals to propagate at a constant speed.

### Distances in special relativity

For a bit of a simpler case, let's consider distances in special relativity first, where the metric is everywhere flat as

\begin{equation} ds^2 = -dt^2 + dx^2 \end{equation}The obvious theoretical distance between two objects, given the canonical foliation, is

\begin{equation} d(\gamma_1, \gamma_2) = \| x_2(t) - x_1(t) \| \end{equation}This is entirely well-defined (although requires the gauge $t(\tau) = \tau$ here and depends on the foliation), but it is not measurable directly. Instead what we need to have is an exchange of signals, so that every process will remain local.

The first observer $\gamma_1$ emits a light signal $\ell_1$ at $\gamma_1(\tau_1)$, which reaches the observer $\gamma_2$ at $\gamma_2(\tau_2)$ before re-emitting it towards $\gamma_1$, which receives it at $\gamma_1(\tau_3)$.

From the point of view of $\gamma_1$, the number of informations we can extract from the light signal are somewhat limited.

- We know (up to some uncertainty) the proper time it took the light signal to bounce back, $\tau_3 - \tau_1$
- If properly encoded, the light signal can encode the proper time at which the signal was received by $\gamma_2$, $\tau_2$, although relating it to $\gamma_1$'s own proper time is delicate.
- Given a standard of the emission, we can deduce the redshift of the signal, giving us the speed (and, for a sufficiently long signal emitted, the acceleration) that both observers are relative to one another (ie the speed of $\gamma_1$ and $\gamma_2$ affect the redshift of the first signal, and the same is true for the return signal).

The common way to measure the distance of two observers with this method is the *Synge formula* :

...

## 3. Measuring angles in general relativity

Comparatively to times and distances, angles are fairly easy to do as they are very much a local measurement.

## 4. Determination of the metric tensor

Once we have all these elements, all that remains for us is to attempt to determine the components of the metric tensor in the coordinates we have defined for our observer. If we have the full measurements of every observer in our grid, this should be doable.

## Measurements in the ADM formalism

For later use, it will be practical to work out some of those measurement issues in the ADM formalism, where we have the decomposition of our metric as

\begin{equation} ds^2 = -N^2 dt^2 + \gamma_{ij} (dx^i + \beta^i dt) (dx^j + \beta^j dt) \end{equation}So that

\begin{eqnarray} g_{00} &=& -N^2 + \beta_k \beta^k\\ g_{0i} &=& \beta_i\\ g_{ij} &=& \gamma_{ij} \end{eqnarray}with the three functions $N$, $\beta$ and $\gamma$. $N$, the lapse function, and $\beta$, the shift function

What we want is to have $N$ and $\beta$ to correspond to our family of observers. In the formalism, our observers $\gamma$ map

\begin{eqnarray} \gamma : \mathbb{I} &\to& \mathbb{R} \times \mathbb{R}^3 \\ \tau &\mapsto& (t(\tau), x^i(\tau)) \end{eqnarray}If we want the time coordinate to be such that $\Delta t = \Delta \tau$ for our curves, that will be

\begin{eqnarray} \Delta \tau = \int_{\tau_1}^{\tau_2} \sqrt{- g(\gamma', \gamma')} d\tau\\ &=& \int_{\tau_1}^{\tau_2} \sqrt{- g(\gamma', \gamma')} d\tau \end{eqnarray}...

This implies that

\begin{equation} \beta^k \beta_k = N^2 \end{equation}...

\begin{eqnarray} a &=& \ddot{\gamma}\\ &=& \nabla_\dot{\gamma} \dot{\gamma} \\ &=& u^\mu (\partial_\mu u^\nu + {\Gamma^\nu_{\mu\sigma}} u^\sigma) \end{eqnarray}## The test field limit, approximations and far-off sources

There are a number of approximations that are essential to experimental general relativity :

- The various objects we use for measurements do not contribute significantly to the gravitational field we want to study. This is the
*test field limit*. - If we can approximate the sources to some degree of error, the deviations from this approximation will have negligible effects on the gravitational field.
- The influence of very far-off sources can be neglected

All three of those assumptions basically boil down to an estimation problem of the Einstein field equation.

### The test field limit

In the test field limit, we presume that the real stress-energy tensor differs from the one we're using by a small amount in a compact region (the region isn't necessarily compact of course, as a charge will for instance have influences reaching potentially to infinity, but such considerations will be left for the following approximations). In other words, we have

\begin{equation} T_{R} = T_{A} + \Pi_S \Delta T \end{equation}Inside a top-hat function $\Pi$ of a region (generally small compared to the system we study), there is a stress-energy tensor which is small in some sense. This will generally be small compared to the source we're considering.

### Source approximation

The problem of the source approximation is fairly close to the test field limit problem, but the

### Error estimation

\begin{equation} G[g + \Delta g] = \kappa (T + \Delta T) \end{equation}Things are not simplified by the appearance of the non-linear in $g$ Einstein tensor, so for now, let's consider the linearized gravity equation in the de Donder gauge

\begin{equation} \Box (h + \Delta h) = \kappa (T + \Delta T) \end{equation}Here $\Delta h$ is the error on the calculated deviation from Minkowski space.

\begin{equation} \Box (h + \Delta h) - \Box (h) = \kappa (T[h + \Delta h] + \Delta T[h + \Delta h]) - \kappa (T[h]) \end{equation}For most sources of matter, $T$ is generally linear in the metric [CHECK WRT INVERSE METRIC], so that we can use

\begin{equation} \Box (\Delta h) = \kappa (T[\Delta h] + \Delta T[h + \Delta h]) \end{equation}Green's function :

\begin{equation} \Delta h = \kappa \int G(x, s) \left[(T[\Delta h](s) + \Delta T[h](s) + \Delta T[\Delta h](s)) \right] ds \end{equation}...

ADM :

As we want to predict the evolution of a system from some initial values, the ADM formalism is advised here.

\begin{eqnarray} \dot{\gamma_{ij}} - \mathcal{L}_\beta \gamma_{ij} &=& -2N K_{ij}\\ \dot{K_{ij}} - \mathcal{L}_\beta K_{ij} &=& -D_i D_j N + N \left[ R_{ij} + K K_{ij} - 2 K_{ik} {K^k}_j + 4 \pi ((S - E) \gamma_ij - 2 S_{ij}) \right]\\ R + K^2 - K_{ij} K^{ij} &=& 16\pi E\\ D_j {K^j}_i - D_i K &=& 8 \pi p_i \end{eqnarray}Now let's suppose that, given some initial conditions $(\Sigma_0, N, \beta, \gamma, K, E, p_i, S)$, our measured values are, for actual dynamical variables, $(\bar{\gamma}, \bar{K}, \bar{E}, \bar{p_i}, \bar{S})$, with for each variable the relation $q = \bar{q} + \delta q$. We also know the behaviour of $E, p_i, S$ given by some equation of motion of the individual fields or an equation of state. What is the error of the prediction on those measured values, given some bounds on $\delta q$?

...

Our computed solution $\bar{\gamma}$ is such that its restriction on $\Sigma_0$ gives us our initial conditions, and it obeys the ADM equations. Our real solution will be of the form

\begin{eqnarray} \dot{\bar{\gamma}}_{ij} + \delta\dot{\gamma}_{ij} - \mathcal{L}_\beta \bar{\gamma}_{ij} - \mathcal{L}_\beta \delta \gamma_{ij} &=& -2N K_{ij}\\ ----\dot{K_{ij}} - \mathcal{L}_\beta K_{ij} &=& -D_i D_j N + N \left[ R_{ij} + K K_{ij} - 2 K_{ik} {K^k}_j + 4 \pi ((S - E) \gamma_ij - 2 S_{ij}) \right]\\ ----R + K^2 - K_{ij} K^{ij} &=& 16\pi E\\ ----D_j {K^j}_i - D_i K &=& 8 \pi p_i \end{eqnarray}We want to know what the measurement at $\gamma(t + \Delta t)$ will be.

## Real measurements of the spacetime metric

Now that we have a theory for the measurement in our extremely idealized case, let's consider a more realistic scenario.

The first obvious limitation we have are the number of measuring apparatus and their precision. Instead of our apparatus foliating the spacetime, we have a finite number $N$ of them, each one with some index $n \in [1 \ldots N]$. As we do not know the geometry of spacetime quite yet, it's a bit early to talk about our measuring apparatus on a grid of appropriate spacing, although this may be done later on. From what we've seen on clocks, there is also only so much we can do for the precision of our clocks. Not only that, but as we are encoding data in those light signals, those are not instantaneous either : each train of data takes a certain time to be emitted, as well as being decoded

What we are basically getting from all our data (ignoring for now the other uncertainties related to instrumentation) is a triangulation of our spacetime manifold into a simplicial complex, similar to the one we'd get in Regge calculus. If we had for instance two observers in Minkowski space exchanging one light signal, from the point of view of the emitting observer, the simplicial complex would be comprised of four points : $t_1$, the sending of our light signal, $t_2$, the receiving of the response, $t$ the bounding from the other observer and $t'$ the synchronized point on its own trajectory. There are a total of $5$ edges here : $t_1t$ and $tt_2$ are both null paths of lengths $0$, $t_1t'$ and $t'2$ are both timelike and of known proper time, and finally $t't$ is spacelike and of known distance.

A metric is essentially infinite-dimensional and here our measurements are not even directly related directly related to it. So first, let's consider the equivalence class involved here. We'll say that two metrics $g_1$, $g_2$ fit the measurements if the resulting spacetime can be skeletonized

### Injectivity radius in spacetimes

One of our big obstacle for measurement is the cut locus. Without knowing how large a normal neighbourhood around a point is, we won't be able to determine what's an appropriate scale for our measurements. Minkowski space's normal neighbourhoods may cover the entire spacetime, while some kind of Wheeler-style quantum foam could have one below experimental reach, given small enough wormholes riddling the spacetime.

What we need is some bounds on the injectivity radius of the exponential map linked to reasonably measurable quantities. This will not be entirely satisfying since any measurement we can perform will have to rely on the spacetime geometry, giving us some kind of bootstrapping problem, and as mentionned, we could have some quantum foam situation, where the very reasonable measurable quantities are just averages of extremely variable densities. But we need to draw the line somewhere, and for now we'll assume that matter, at least in our vicinity, has some reasonable bounds on the stress energy tensor.

There are many ways to deal with the injectivity radius in spacetimes, as we can't use directly the Riemannian definition, lacking a distance function.

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Last updated :

*2020-01-13 14:57:35*