191 lines
8.0 KiB
TeX
191 lines
8.0 KiB
TeX
\appendix
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\section{\large Gravitational Waves in Linearised General Relativity}\label{sec:appendix}
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In general relativity we consider pseudo-Riemannian manifolds in which
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the interval is given by (assuming the summation convention)
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\begin{equation}
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ds^2=g_{\mu\nu}(x)dx^{\mu}dx^{\nu}
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\end{equation}
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where $g_{\mu\nu}$ are the components of the metric tensor field in
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the chosen coordinate system. A weak gravitational field corresponds
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to a region of spacetime that is only slightly curved. Thus, there
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exist coordinate systems $x^{\mu}$ in which the spacetime metric takes
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the form
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\begin{equation}\label{eq:linear}
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g_{\mu\nu} = \eta_{\mu\nu} + h_{\mu\nu}
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\end{equation}
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where $|h_{\mu\nu}| \ll 1$, the partial derivatives of $h_{\mu\nu}$
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are also small and $\eta_{\mu\nu}$ are the components of the metric
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tensor in flat Minkowski spacetime.
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It can be shown that for infinitesimal general coordinate
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transformations of the form
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\begin{equation}
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x^{\prime\mu} = x^{\mu}+\xi^{\mu}(x),
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\end{equation}
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working to first order in small quantities, the metric transforms as
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follows:
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\begin{equation}
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g^{\prime}_{\mu\nu} = \eta_{\mu\nu} + h_{\mu\nu} - \partial_\mu\xi_\nu
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- \partial_\nu\xi_\mu.
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\end{equation}
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We note that $g^{\prime}_{\mu\nu}$ is of the same form
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as \eqref{eq:linear}, hence the new metric perturbation functions are
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related to the old ones via
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\begin{equation}\label{eq:gauge}
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h^{\prime}_{\mu\nu} = h_{\mu\nu} - \partial_\mu\xi_\nu
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- \partial_\nu\xi_\mu.
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\end{equation}
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We consider $h_{\mu\nu}$ to be a tensor field defined on the flat
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Minkowski background spacetime, hence \eqref{eq:gauge} can be
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considered as analogous to a gauge transformation in
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electromagnetism. If $h_{\mu\nu}$ is a solution to the linearised
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gravitational field equations then the same
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physical situation is desribed by \eqref{eq:gauge}. However, in this
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viewpoint \eqref{eq:gauge} is viewed as a gauge transformation rather
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than a coordinate transformation. We are still working in the same set
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of coordinates $x^\mu$ and have defined a new tensor whose components
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in this coordinate system are given by \eqref{eq:gauge}. Here we adopt
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the viewpoint that $h_{\mu\nu}$ is a symmetric tensor field (under
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global Lorentz transformations) defined in quasi-Cartesian coordinates
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on a flat Minkowski background spacetime.
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The linearised field equations of general relativity can be written as
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\begin{equation}
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\Box^2\bar{h}^{\mu\nu} = -2\kappa T^{\mu\nu},
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\end{equation}
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where $T^{\mu\nu}$ is the energy-momentum tensor,
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$\bar{h}^{\mu\nu} \equiv h^{\mu\nu} - \frac{1}{2}\eta^{\mu\nu}h$, $h =
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h^\mu_\mu$, provided that the $\bar{h}^{\mu\nu}$ components satisfy
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the Lorenz gauge condition
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\begin{equation}
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\partial_\mu\bar{h}^{\mu\nu} = 0.
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\end{equation}
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In vacuo, where $T^{\mu\nu} = 0$, the general solution of the
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linearised field equations may be written as a superposition of
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plane-wave solutions of the form
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\begin{equation}\label{eq:wave}
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\bar{h}^{\mu\nu} = A^{\mu\nu}\exp(ik_\rho x^\rho),
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\end{equation}
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where $A^{\mu\nu}$ are constant components of a symmetric tensor and
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$k_\mu$ are the constant, real components of a vector. The Lorenz gauge
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condition is satisfied provided
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\begin{equation}
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A^{\mu\nu}k_\nu = 0.
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\end{equation}
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Physical solutions may be obtained by taking the real part
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of \eqref{eq:wave}.
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Further gauge transformations will preserve the Lorenz gauge condition
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provided that the four functions $\xi^\mu(x)$ satisfy $\Box^2\xi^\mu =
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0$. A common choice for plane gravitational waves is the
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transverse-traceless gauge defined by choosing
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\begin{equation}\label{eq:TT}
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\bar{h}^{0i}_{TT} \equiv 0 \,\,\,\,\,\, \text{and} \,\,\,\,\,\, \bar{h}_{TT} \equiv 0,
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\end{equation}
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where latin alphabet indices run over the spatial dimensions
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only. Furthermore the Lorenz gauge condition gives the constraints
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\begin{equation}\label{eq:lorenzTT}
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\partial_0\bar{h}^{00}_{TT} = 0 \,\,\,\,\,\, \text{and} \,\,\,\,\,\, \partial_i\bar{h}^{ij}_{TT} = 0.
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\end{equation}
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We note that the first condition in \eqref{eq:lorenzTT} implies that
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for non-stationary perturbations $\bar{h}^{00}_{TT}$ also
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vanishes. Therefore, in this gauge only the spatial components
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$\bar{h}^{ij}_{TT}$ are non-zero.
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For a particular case of an arbitrary plane gravitational wave of the
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form \eqref{eq:wave} the conditions \eqref{eq:TT} imply that
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\begin{equation}
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A^{0i}_{TT} = 0 \,\,\,\,\,\, \text{and} \,\,\,\,\,\, (A_{TT})^\mu_\mu = 0
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\end{equation}
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and the Lorenz gauge conditions in \eqref{eq:lorenzTT} require that
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\begin{equation}
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A^{00}_{TT} = 0 \,\,\,\,\,\, \text{and} \,\,\,\,\,\, A^{ij}_{TT}k_j = 0.
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\end{equation}
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Under these conditions the $A^{\mu\nu}_{TT}$ components are given by
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\begin{equation}
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A^{ij}_{TT} = (P^i_kP^j_l - \frac{1}{2}P^{ij}P_{kl})A^{kl},
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\end{equation}
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where $P_{ij} \equiv \delta_{ij} - n_in_j$ and $n_i = \hat{k}_i$ \cite{hobson}.
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\section{The Algorithm}
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We solve the two-particle Schr\"{o}dinger equation for the Hamiltonian
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\eqref{eq:main} via the explicit staggered method \cite{numerics}. We
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choose our units such that $\hbar = 1$ and the neutron mass $m_n =
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1$. The wave function is evaluated on a grid of discrete values for
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the independent variables
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\begin{equation}
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\Psi(x_1, x_2, t) = \Psi(x_1 = l\Delta x, x_2 = m\Delta x, t = n\Delta
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t) \equiv \Psi^n_{l, m},
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\end{equation}
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where $l$, $m$ and $n$ are integers. We separate it into real and
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imaginary parts,
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\begin{equation}
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\Psi^n_{l,m} = u^{n+1}_{l,m} + iv^{n+1}_{l,m},
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\end{equation}
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which allows us to solve the Schr\"{o}dinger equation via a finite
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difference method with a pair of coupled equations:
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\begin{samepage}
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\begin{equation}
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u^{n+1}_{l,m} = u^{n-1}_{l,m} - \left\{ \alpha \left[ v^n_{l+1,m} +
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v^n_{l-1,m} + v^n_{l,m+1} + v^n_{l,m-1} - 4v^n_{l,m} \right] - 2
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\Delta t V_{l,m}v^n_{l,m} \right\} ,
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\end{equation}
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\begin{equation}
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v^{n+1}_{l,m} = v^{n-1}_{l,m} + \left\{ \alpha \left[ u^n_{l+1,m} +
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u^n_{l-1,m} + u^n_{l,m+1} + u^n_{l,m-1} - 4u^n_{l,m} \right] - 2
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\Delta t V_{l,m}u^n_{l,m} \right\},
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\end{equation}
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\end{samepage}
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where
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\begin{equation}
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\alpha = g^{xx}\frac{\Delta t}{\Delta x^2},
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\end{equation}
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and $V_{l,m}$ is the combined value of the potential due to the
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harmonic well and neutron-neutron interaction on the discretised
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grid given by
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\begin{equation}
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V_{l,m} = \frac{1}{2}\omega^2(l^2 + m^2)\Delta x^2 + \nu\delta_{l,m}.
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\end{equation}
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In order to evaluate the value of the real part at step $n+1$ we need
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to know the values of the real part at $n-1$ and the imaginary part at
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$n$. The same is true for evaluating the imaginary part. Therefore, we
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do not have to calculate both parts at every time step and we compute
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the real and imaginary parts at slightly different (staggered)
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times. The form of the discretised equations is identical to those
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presentied in \cite{numerics}. However, the value of $\alpha$ is now
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scaled by the oscillating metric tensor component $g^{xx}$. This novel
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modification does not lead to instabilities and simulations have been
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confirmed to conserve probability to the same precision as before.
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For the more general Laplacian \eqref{eq:general}, the equations have to
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be modified even further and new terms are introduced for the
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wavefunction's first derivative terms present
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\begin{equation}
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\begin{array}{lcl}
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u^{n+1}_{l,m} & = & u^{n-1}_{l,m} - \left\{ \alpha \left[ v^n_{l+1,m}
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+ v^n_{l-1,m} + v^n_{l,m+1} + v^n_{l,m-1} -
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4v^n_{l,m} \right] \right. \\\\ & & \left. + \beta \left[ v^n_{l+1,m} -
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v^n_{l-1,m} + v^n_{l, m+1} - v^n_{l, m-1} \right] - 2
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\Delta t V_{l,m}v^n_{l,m} \right\} ,
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\end{array}
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\end{equation}
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\begin{equation}
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\begin{array}{lcl}
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v^{n+1}_{l,m} & = & v^{n-1}_{l,m} + \left\{ \alpha \left[ u^n_{l+1,m}
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+ u^n_{l-1,m} + u^n_{l,m+1} + u^n_{l,m-1} -
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4u^n_{l,m} \right] \right. \\\\ & & \left. + \beta \left[
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u^n_{l+1,m} - u^n_{l-1,m} + u^n_{l, m+1} - u^n_{l, m-1} \right] - 2
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\Delta t V_{l,m}u^n_{l,m} \right\},
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\end{array}
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\end{equation}
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where
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\begin{equation}
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\beta = \frac{\partial g^{xx}}{\partial x} \frac{\Delta t}{2 \Delta x}.
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\end{equation}
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The two components can still be evaluated at staggered times. This
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further modification does not cause instabilities and conserves
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probability very well.
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